Matrix Bases for Star Products: a Review
We review the matrix bases for a family of noncommutative ⋆ products based on a Weyl map. These products include the Moyal product, as well as the Wick-Voros products and other translation invariant ones. We also review the derivation of Lie algebra type star products, with adapted matrix bases. We...
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| Цитувати: | Matrix Bases for Star Products: a Review / F. Lizzi, P. Vitale // Symmetry, Integrability and Geometry: Methods and Applications. — 2014. — Т. 10. — Бібліогр.: 83 назв. — англ. |
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irk-123456789-1466142019-02-11T01:23:13Z Matrix Bases for Star Products: a Review Lizzi, F. Vitale, P. We review the matrix bases for a family of noncommutative ⋆ products based on a Weyl map. These products include the Moyal product, as well as the Wick-Voros products and other translation invariant ones. We also review the derivation of Lie algebra type star products, with adapted matrix bases. We discuss the uses of these matrix bases for field theory, fuzzy spaces and emergent gravity. 2014 Article Matrix Bases for Star Products: a Review / F. Lizzi, P. Vitale // Symmetry, Integrability and Geometry: Methods and Applications. — 2014. — Т. 10. — Бібліогр.: 83 назв. — англ. 1815-0659 2010 Mathematics Subject Classification: 58Bxx; 40C05; 46L65 DOI:10.3842/SIGMA.2014.086 http://dspace.nbuv.gov.ua/handle/123456789/146614 en Symmetry, Integrability and Geometry: Methods and Applications Інститут математики НАН України |
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We review the matrix bases for a family of noncommutative ⋆ products based on a Weyl map. These products include the Moyal product, as well as the Wick-Voros products and other translation invariant ones. We also review the derivation of Lie algebra type star products, with adapted matrix bases. We discuss the uses of these matrix bases for field theory, fuzzy spaces and emergent gravity. |
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Lizzi, F. Vitale, P. |
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Lizzi, F. Vitale, P. Matrix Bases for Star Products: a Review Symmetry, Integrability and Geometry: Methods and Applications |
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Lizzi, F. Vitale, P. |
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Lizzi, F. |
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Matrix Bases for Star Products: a Review |
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Matrix Bases for Star Products: a Review |
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Matrix Bases for Star Products: a Review |
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Matrix Bases for Star Products: a Review |
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Matrix Bases for Star Products: a Review |
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matrix bases for star products: a review |
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Інститут математики НАН України |
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2014 |
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Matrix Bases for Star Products: a Review / F. Lizzi, P. Vitale // Symmetry, Integrability and Geometry: Methods and Applications. — 2014. — Т. 10. — Бібліогр.: 83 назв. — англ. |
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Symmetry, Integrability and Geometry: Methods and Applications |
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AT lizzif matrixbasesforstarproductsareview AT vitalep matrixbasesforstarproductsareview |
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2025-07-11T00:19:51Z |
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Symmetry, Integrability and Geometry: Methods and Applications SIGMA 10 (2014), 086, 36 pages
Matrix Bases for Star Products: a Review?
Fedele LIZZI †‡§ and Patrizia VITALE †‡
† Dipartimento di Fisica, Università di Napoli Federico II, Napoli, Italy
E-mail: fedele.lizzi@na.infn.it, patrizia.vitale@na.infn.it
‡ INFN, Sezione di Napoli, Italy
§ Institut de Ciéncies del Cosmos, Universitat de Barcelona, Catalonia, Spain
Received March 04, 2014, in final form August 11, 2014; Published online August 15, 2014
http://dx.doi.org/10.3842/SIGMA.2014.086
Abstract. We review the matrix bases for a family of noncommutative ? products based
on a Weyl map. These products include the Moyal product, as well as the Wick–Voros
products and other translation invariant ones. We also review the derivation of Lie algebra
type star products, with adapted matrix bases. We discuss the uses of these matrix bases
for field theory, fuzzy spaces and emergent gravity.
Key words: noncommutative geometry; star products; matrix models
2010 Mathematics Subject Classification: 58Bxx; 40C05; 46L65
1 Introduction
Star products were originally introduced [34, 62] in the context of point particle quantization,
they were generalized in [9, 10] to comprise general cases, the physical motivations behind this
was the quantization of phase space. Later star products became a tool to describe possible
noncommutative geometries of spacetime itself (for a review see [4]). In this sense they have
become a tool for the study of field theories on noncommutative spaces [77]. In this short review
we want to concentrate on one particular aspect of the star product, the fact that any action
involving fields which are multiplied with a star product can be seen as a matrix model. This
is because the noncommutative star products algebras can always be represented as operators
on a suitable Hilbert space, and one can then just consider the matrix representation of these
operators. The matrix representation is not only a useful tool, but also offers a conceptual
interpretation of noncommutative spaces. It shows that, in analogy with quantum phase space,
the proper setting of deformed products is noncommutative geometry [20], seen as a spectral
theory of operators.
This is a partial review, as we do not have the ambition to cover exhaustively all known
products, nor all known bases. We will consider products which are connected to symmetries
and which have developments in field theory. Even then we will not cover completely all work
done, and apologize for the omissions.
In order to construct the matrix bases we will use mainly quantization-dequantization maps,
whereby we associate operators to functions, and viceversa. We will first consider a family of
translation-invariant star products induced on suitable algebras of functions on R2n through
s-ordered quantization-dequantization maps. In this approach noncommutative algebras are
obtained as symbols of algebras of operators, which are defined in terms of a special family of
operators using the trace formula (what is sometimes called the ‘dequantization’ map because
of its original meaning in the Wigner–Weyl formalism), while the reconstruction of operators in
?This paper is a contribution to the Special Issue on Deformations of Space-Time and its Symmetries. The
full collection is available at http://www.emis.de/journals/SIGMA/space-time.html
mailto:email@address
mailto:email@address
http://dx.doi.org/10.3842/SIGMA.2014.086
http://www.emis.de/journals/SIGMA/space-time.html
2 F. Lizzi and P. Vitale
terms of their symbols (the ‘quantization’ map) is determined using another family of operators.
These two families determine completely the noncommutative algebra, including the kernel of
the star-product.
Once we have established the connection, in Section 4 we construct matrix bases for s-ordered
products and for the Moyal and Wick–Voros cases in more detail. In Section 5 we review the
construction of a family of star products of Lie algebra type, obtained by reduction of higher-
dimensional, translation invariant ones. We thus discuss in Section 6 the matrix basis for one
of them on the space R3. Section 7 is devoted to the application to quantum field theory of the
matrix bases introduced previously. Finally, Section 8 is devoted to fuzzy spaces.
2 The star product
In this section we review a general construction for the noncommutative algebra of operator
symbols acting on a Hilbert space H. We follow the presentation and the notation of [56, 57].
Given an operator  acting on the Hilbert space H (which can be finite- or infinite-dimen-
sional), let us have two distinct operator bases for the operators acting on H, Û(x) and D̂(x).
The two bases are labelled by a set of parameters x = (x1, . . . , xN ), with xk ∈ K, k = 1, . . . , N ,
with K any field, usually the reals, complex or integers. The symbol fA(x) of the operator  is
the following function of x:
fA = Ω−1
(
Â
)
≡ tr
(
ÂÛ(x)
)
. (2.1)
We assume that the trace exists for all parameters x, although often the function f may be
a distribution. If we now consider the “parameters” x as the coordinates of a manifold it makes
sense to define the inverse map (the “reconstruction formula”), which associates operators to
functions, is defined in terms of the second family of operators, that is
 = Ω̂(fA) =
∫
fA(x)D̂(x)dx. (2.2)
The map Ω̂ is often called the quantization map, because of its role in quantum mechanics. In
the standard quantization scheme, which associates Hermitian operators to real functions, it is
the Weyl map, whereas its inverse is the Wigner map. We assume that an appropriate measure
dx exists to make sense of the reconstruction formula and that the two maps have a sufficiently
rich domain to ensure that the maps are invertible for a dense set in the space of functions.
In this way we have an invertible map between functions on space and operators on a Hilbert
space.
The symbols form an associative algebra endowed with a noncommutative (star) product
defined in terms of the operator product
fA ? fB(x) = Ω−1
(
ÂB̂
)
= tr
(
ÂB̂Û(x)
)
, (2.3)
where the associativity follows from the associativity of the operator product.
The star product may be expressed in terms of an integral kernel
fA ? fB(x) =
∫
fA(x′)fB(x′′)K(x′, x′′, x)dx′dx′′, (2.4)
with
K(x′, x′′, x) = tr
(
D̂(x′)D̂(x′′)Û(x)
)
. (2.5)
Matrix Bases for Star Products: a Review 3
The operators D̂(x) and Û(x) are also known as quantizer and dequantizer. The associativity
condition implies∫
K(x′, y, x)K(x′′, x′′′, y)dy =
∫
K(x′, x′′, y)K(y, x′′′, x)dy.
From the compatibility request of equations (2.1), (2.2) we easily derive an important condition
that the two families of basic operators have to satisfy
tr
(
D̂(x′)Û(x)
)
= δ(x′ − x), (2.6)
where δ is to be replaced with the Kronecker delta for discrete parameters. Going back to (2.5),
it is to be stressed that the kernel of the star product has been obtained solely in terms of the
operators D(x) and U(x), which in turn are only constrained by (2.6). Thus, to each pair of
operators satisfying (2.6) it is associated an associative algebra with a star product. The role
of D(x) and U(x) can be exchanged, what gives rise to a duality symmetry [56]. This duality
allows for the definition of a new star product, different from the one in (2.4) (unless Û and D̂
are proportional), defined in terms of a dual kernel
Kd(x′, x′′, x) = tr
(
Û(x′)Û(x′′)D̂(x)
)
.
3 The Weyl–Wigner and the s-ordered maps
In this section we review the usual Moyal product on the two-dimensional plane R2 and the
family of s-ordered products which generalizes it. The generalization is made in terms of the
quantization-dequantization maps illustrated above.
We first review a few properties of coherent states and their relation to the number basis,
which shall be used in the rest of this section. Let x = (x1, x2) ∈ R2. We use a slightly
unconventional notation for the coordinates on the complex plane defining:
z =
1√
2
(x1 + ix2).
We define the creation and annihilation operators on the plane, a†, a, as
a† =
1√
2
(x̂1 − ix̂2), a =
1√
2
(x̂1 + ix̂2)
with commutation relation
[a, a†] = θ,
where θ is a dimensional parameter of area dimension. Coherent states of the plane are thus
defined by a|z〉 = z|z〉. The decomposition of the identity reads
1 =
1
πθ
∫
d2z|z〉〈z|.
Coherent states are non-orthogonal
〈z|z′〉 = e−
1
θ
(z̄z+z̄′z′−2z̄z′).
Given the number operator N̂ = a†a, and indicating with |n〉 its eigenvalues, we state the useful
relations
a|n〉 =
√
nθ|n− 1〉, a†|n〉 =
√
(n+ 1)θ|n+ 1〉, |n〉 =
(
a†
)n
√
n!θn
|0〉 (3.1)
together with
〈z|n〉 = e−
z̄z
2θ
z̄n√
n!θn
.
4 F. Lizzi and P. Vitale
3.1 The Moyal product
Given an operator  acting in the Hilbert space of square integrable functions on R, the Wigner–
Weyl symbol of  and the reconstruction map are defined by means of the families of operators
ÛM(x) = 2 D̂(z, z̄)(−1̂)a
†aD̂(−z,−z̄), (3.2)
D̂M(x) =
1
(2π)2
ÛM(x), (3.3)
where (−1̂)a
†a is the parity operator. To simplify notations we assume θ = 1 unless otherwise
stated. D̂(z, z̄) is the unitary displacement operator realizing the ray representation of the group
of translations of the plane
D̂(z, z̄) = eza
†−z̄a, D̂(z, z̄)|w〉 = |w + z〉e
1
2
(zw̄−wz̄). (3.4)
We shall use the isomorphism R2 = C in the following and the notation D̂(z) as a shorthand
for D̂(z, z̄) to indicate the functional dependence on the plane coordinates from now on.
The compatibility condition (2.6) between the operators (3.2) and (3.3) is then readily verified
on using well known properties of D̂(z)
(−1̂)a
†aD̂(z)(−1̂)a
†a = D̂(−z), tr D̂(z) = πδ(Re z)δ(Im z). (3.5)
The quantizer and dequantizer may be expressed in the more familiar form
D̂M(x) =
1
(2π)2
ÛM(x) =
∫
d2ξ
(2π)2
exp
[
i
(
ξ1(x̂2 − x2) + ξ2(x̂1 − x1)
)]
. (3.6)
The Weyl quantization map obtained by the reconstruction formula (2.2) therefore reads
Ω̂(f) = f̂ =
∫
d2xf(x)
∫
d2ξ
(2π)2
exp
[
i
(
ξ1(x̂2 − x2) + ξ2(x̂1 − x1)
)]
,
whereas the inverse map obtained by (2.2) is represented by:
f(x) = tr
(
f̂
∫
d2ξ exp
[
i
(
ξ1(x̂2 − x2) + ξ2(x̂1 − x1)
)])
.
The algebra of symbols so defined is endowed with the well known Moyal product. In this
language it is defined in terms of the kernel, which is easily obtained specializing (2.5) to this
case. On restoring the noncommutative parameter θ and using the first of equations (3.5)
together with the product rule for the displacement operators
D̂(z)D̂(z′) = D̂(z + z′)ei Im(zz̄′)
we have indeed
KM(x′, x′′, x) = C exp
{
2iεij
[
xix
′
j + x′ix
′′
j + x′′i xj
]}
(3.7)
with C a normalization constant and εij the antisymmetric tensor. Its expression in complex
coordinates is also useful for future comparison
KM(z′, z′′, z) = C exp
{
4i Im
[
z̄′z′′ + zz̄′′ + z′z̄
]}
. (3.8)
The Moyal product is then defined as
(f ?M g)(x) =
∫
d2x′d2x′′f(x′)g(x′′)KM(x′, x′′, x). (3.9)
Matrix Bases for Star Products: a Review 5
In complex coordinates the relation (3.9) reads
(f ?M g)(z, z̄) =
∫
d2z′d2z′′f(z′, z̄′)g(z′′, z̄′′)e{4i Im[z̄′z′′+zz̄′′+z′z̄]}.
In the case of the coordinate functions on R2, upon restoring the factors of θ, this gives the
usual star-commutation relation
[x1, x2]?M = iθ (3.10)
or equivalently
[z, z̄]?M = θ. (3.11)
The Moyal product can be given the popular form
(f ?M g)(z, z̄) = f(z, z̄) exp
[
θ
2
(←−
∂z
−→
∂z̄ −
←−
∂z̄
−→
∂z
)]
g(z, z̄), (3.12)
where the operator
←−
∂ (resp.
−→
∂ ) acts on the left (resp. on the right). It is well known that this
form is obtained by the integral expression through an asymptotic expansion in the parame-
ter θ [25].
The generalization to the higher-dimensional cases is straightforward.
What is properly defined as the Moyal algebra isMθ :=ML(R2
θ)∩MR(R2
θ) whereML(R2
θ),
the left multiplier algebra, is defined as the subspace of tempered distributions that give rise
to Schwartz functions when left multiplied by Schwartz functions; the right multiplier alge-
bra MR(R2
θ) is analogously defined. For more details we refer to the appendix in [32] and refe-
rences therein. In the present article we can think of Mθ either as the algebra of ?-polynomial
functions in z, z̄ properly completed, or the algebra generated by Schwarz functions, or a variant.
One should however pay attention since the domains of definition depend on the particular form
of the product. Its commutative limit, F(R2), is the commutative multiplier algebra OM(R2),
the algebra of smooth functions of polynomial growth on R2 in all derivatives [29].
3.2 s-ordered symbols
s-ordered symbols have been introduced long ago by Cahill and Glauber [16] in the context of
quantum mechanics to deal with different orderings w.r.t. the Weyl–Wigner symmetric ordering
of operators. The corresponding quantization and dequantization maps are often referred to as
weighted maps. We refer to [57] and references therein for more details. Here the stress will be
on the different star-products which are related to such orderings. We consider the two families
of operators Ûs(x) and D̂s(x), x ∈ R2, of the form [16]
Ûs(x) =
2
(1 + s)
D̂(z)
(
s− 1
s+ 1
)a†a
D̂(−z) =
2
1 + s
(
s− 1
s+ 1
)(a†−z̄)(a−z)
, (3.13)
D̂s(x) =
1
(2π)2
Û−s(x), (3.14)
where s is chosen to be real in the analysis below. We shall see that the convergence of the kernel
will impose constraints on the allowed values of s. The case s = 0 corresponds to the standard
Wigner–Weyl situation described above. Two other relevant cases correspond to the singular
limits s = ±1 which yield respectively normal and anti-normal ordering in quantization. It is
interesting to notice that these two cases are in the duality relation discussed previously (namely
6 F. Lizzi and P. Vitale
the role of Û and D̂ are exchanged). The duality symmetry then connects normal and anti-
normal ordering. The Wigner–Weyl Moyal quantization scheme selects instead the symmetric
ordering, consistently with it being self-dual.
For s 6= ±1 the star-product kernel of s-ordered symbols is calculated along the same lines
as the Moyal kernel (3.7). To our knowledge it was first calculated in [56], although no analysis
on the convergence was made. Let us sketch the derivation. For simplicity we define q =
(s + 1)/(s − 1), q̃ = q−1 and we skip overall constants. From the definition (2.5) and the
dequantizer and quantizer operators (3.13), (3.14), we have
Ks(z
′, z′′, z) = tr
(
D̂(z̄′, z′)D̂(z̄′′, z′′)Û(z̄, z)
)
= (1− q)2(1− q̃)
× tr
(
D̂(−z̄,−z)D̂(z̄′, z′)qa
†aD̂(−z̄′,−z′)D̂(z̄′′, z′′)qa
†aD̂(−z̄′′,−z′′)D̂(z̄, z)q̃a
†a
)
.
Let us introduce
x = −z + z′, y = −z′ + z′′ so that − z′′ + z = −x− y. (3.15)
On using the composition rule for the generalized displacement operator D̂(w̄, z) = exp(za†−w̄a)
D̂(w̄, z)D̂(ū, y) = D̂(w̄ + ū, z + y) exp
(
1
2(zū− yw̄)
)
and the commutation relation
qa
†aD̂(z̄, z) = D̂(q̃z̄, qz)qa
†a,
we arrive at
Ks(z
′, z′′, z) = Ks(x, y) = (1− q)(1− q̃)2 tr
[
D̂((1− q̃)x̄, (1− q)x)qa
†a
]
× exp
[
(1− q)x̄y − (1− q̃)ȳx− 1
2(q − q̃)x̄x
]
. (3.16)
On using the coherent states basis to compute the trace, we have
qa
†a|w〉 = e(q2−1) w̄w
2 |qw〉,
so that we are left with a Gaussian integral
tr
[
D̂((1− q̃)x̄, (1− q)x)qa
†a
]
= e(q̃−q)x̄x
∫
d2we−(1−q)(w̄−x̄)(w−x) =
e(q̃−q)x̄x
1− q
.
On replacing into equation (3.16) we finally have
Ks(x, y) = C(1− q̃)(1− q)e(q̃−q)x̄x+(1−q)x̄y−(1−q̃)ȳx (3.17)
with C an overall constant. This expression can be seen to reproduce the Moyal kernel (3.8) for
s = 0 (q = q̃ = −1). On replacing the original variables z, z′, z′′, it can be seen to coincide with
the result of [56], up to the exchange of q with q̃. We have indeed
Ks(z
′, z′′, z) = C(1− q̃)(1− q)exp
[
(q̃ − q)z̄z + (q̃ − 1)z′z̄′′ + (1− q)z′′z̄′ + (q − 1)z′′z̄
+ (1− q̃)zz̄′′ + (q − 1)zz̄′ + (1− q̃)z′z̄
]
. (3.18)
Let us notice that this expression is singular for s→ ±1. These two cases have to be considered
separately (in the following we will explicitly compute the kernel for the case s = 1). We will
denote with ?s the corresponding star product and with As the noncommutative algebra of
functions on the plane with ?s-noncommutativity, so to have
(f ?s g)(z̄, z) =
∫
d2z′d2z′′f(z̄′, z′)g(z̄′′, z′′)Ks(z
′, z′′, z).
Matrix Bases for Star Products: a Review 7
The convergence of such an expression might impose severe constraints on the algebra of func-
tions selected and has to be analyzed case by case. A careful analysis of the convergence of the
kernel shows that its integral is finite and equal to 1 (as it should) only for −1 < s ≤ 0. Indeed
the introduction of the variables x and y makes it clear that for such a choice of s the integral
of K(x, y) is Gaussian.
As for series expressions of the product, analogues of (3.12) it should be possible to obtain
them on restoring the noncommutative parameter and expanding in terms of it, as in the Moyal
case. We are not aware that such expressions have been investigated before, but we see no
a priori obstruction to perform the calculation.
To illustrate the usefulness of equation (3.17) let us compute the star product of z with z̄ in
detail. We have
z ?s z̄ =
∫
d2z′d2z′′z′z̄′′Ks(z
′, z′′, z).
On expressing z′, z′′ in terms of the variables x and y introduced in (3.15) and replacing the
integral kernel we arrive at
z ?s z̄ = (1− q̃)(1− q)
∫
d2xd2y(x+ z)(x̄+ ȳ + z̄)e(q̃−q)x̄x+(1−q)x̄y−(1−q̃)ȳx,
which may be reduced to a sum of Gaussian integrals. By direct calculation we see that the
integrals containing x̄x, x̄z, z̄x and ȳz are zero. We are left with
z ?s z̄ = (1− q̃)(1− q)
∫
d2xd2y(z̄z + xȳ)e(q̃−q)x̄x+(1−q)x̄y−(1−q̃)ȳx. (3.19)
The first integral is exactly the integral of the kernel, multiplied by the factor z̄z∫
d2xd2ye(q̃−q)x̄x+(1−q)x̄y−(1−q̃)ȳx =
1
(1− q)(1− q̃)
. (3.20)
The second one can be seen to yield∫
d2xd2yȳxe(q̃−q)x̄x+(1−q)x̄y−(1−q̃)ȳx =
1
(1− q)(1− q̃)2
. (3.21)
Thus, replacing (3.20) and (3.21) into (3.19), we obtain the interesting result
z ?s z̄ = z̄z +
1
1− q̃
.
Analogously we obtain
z̄ ?s z = z̄z − 1
1− q
so to obtain for the star commutator
z ?s z̄ − z̄ ?s z = 1,
which is independent of q as expected (we recall that all translation-invariant products are
equivalent from the point of view of the star commutator, being a quantization of the same
Poisson bracket).
In this section we have chosen for simplicity to stick to s, hence q, q̃, real. The analysis can
be repeated for s complex. The convergence of the kernel and of the star product will impose
constraints on the real part of s.
8 F. Lizzi and P. Vitale
3.3 The Wick–Voros product
Let us consider the normal ordered case (s = 1) in more detail. It has been shown [16] that in
such case the limit for the quantizer is singular (its eigenvalues are infinite for all values of z)
and we may represent it as the normal-ordered operator delta function
D̂W(z) =
∫
d2ξeξ(a
†−z̄)e−ξ̄(a−z) (3.22)
with ξ = ξ1 + iξ2, which differs from (3.6) by a phase. Operators in such a scheme acquire the
normal-ordered expression
f̂ =
∞∑
n,m=0
fnm
(
a†
)n
am.
The dequantizer instead is simply the projection operator over coherent states
ÛW(z) = |z〉〈z|,
so that the Wick–Voros star product which is associated to this quantization-dequantization
scheme, can be easily computed by means of equation (2.3) in terms of the expectation value
over coherent states
(f ?W g)(z, z̄) = tr
(
|z〉〈z|f̂ ĝ|
)
= 〈z|f̂ ĝ|z〉 =
∑
n,m,p,q
fnmgpq〈z|
(
a†
)n
am
(
a†
)p
aq|z〉
=
∫
d2ξ
∑
n,m,p,q
fnmgpq z̄
nzq〈z|am|ξ〉〈ξ|
(
a†
)p|z〉
=
∫
d2ξ
∑
n,m,p,q
fnmgpq z̄
nzqξmξ̄p〈z|ξ〉〈ξ|
(
a†
)p|z〉
=
∫
d2ξf(z̄, ξ)g(ξ̄, z)|〈z|ξ〉|2. (3.23)
Alternatively, on using equation (2.5), we may compute the kernel of the star product
KW(z′, z′′, z) = tr
[
D̂(z̄′, z′)D̂(z̄′′, z′′)Û(z̄, z)
]
= 〈z|D̂(z̄′, z′)D̂(z̄′′, z′′)|z〉
=
∫
d2ξ〈z|D̂(z̄′, z′)|ξ〉〈ξ|D̂(z̄′′, z′′)|z〉
and replace the latter in the star product expression
(f ?W g)(z̄, z) =
∫
d2z′d2z′′f(z̄′, z′)g(z̄′′, z′′)KW(z′, z′′, z). (3.24)
On using for example the matrix basis expansion of operators equation (3.3) and the quantization
map equation (2.2) we then observe that
f(z̄, ξ)〈z|ξ〉 = 〈z|f̂ |ξ〉 =
∫
d2z′f(z̄′, z′)〈z|D̂(z̄′, z′)|ξ〉,
which shows that equation (3.24) coincides with the direct computation obtained in (3.23).
As in the case of the Moyal product, upon restoring the parameter θ we can perform a series
expansion, yielding the popular expression
(f ?W g)(z, z̄) = f(z, z̄) exp
[
θ
←−
∂z
−→
∂z̄
]
g(z, z̄). (3.25)
Matrix Bases for Star Products: a Review 9
The Wick–Voros product has been employed in QFT to discuss the emergence of the mixing
independently from the specific translation invariant product chosen [27]. In [8] the Wick–Voros
star product is singled out as it allows for a consistent definition of quantum state.
A word of caution is in order, concerning the domain and the range of the weighted Weyl
map associated to the Wick–Voros quantizer (3.22). While the standard Weyl map associates
to Schwarzian functions Hilbert Schmidt operators, for the weighted Weyl map determined the
Wick–Voros quantizer (3.22) this is not always the case. Explicit counterexamples are discussed
in [48, 49, 50, 51]. The exact correspondence between the appropriate subalgebras of smooth
functions on the plane and bounded operators is discussed in [73] for all s-ordered quantization
schemes, whereas the convergence of the series expansion in (3.25) has been extensively discussed
and established in [11] and references therein.
One important aspect of these products is the fact that derivatives are inner automorphisms:
∂
∂xi
f = θ−1εij [x
j , f ]?
and
∂zf = θ−1[f, z̄]?, ∂z̄f = θ−1[f, z]?. (3.26)
Note that these relations are valid for all products considered in this section. This is a straight-
forward consequence of the fact that the ? commutator (3.11) holds not only for the Moyal
product, but for all s-ordered products.
3.4 Translation invariance
Defining the translation in the plane R2 by a vector a as Ta(f)(x) = f(x + a), by translation
invariant product we mean the property
Ta(f) ? Ta(g) = Ta(f ? g). (3.27)
It is interesting to notice that s-ordered star products described in the previous subsection are
translation invariant. This is an almost obvious consequence of the fact that s-ordered star prod-
ucts are defined in terms of the displacement operator (3.4), which realizes a representation of
the group of translations of the plane. It is however straightforward to check the relation (3.27),
once we observe that the integral kernel for s-ordered products (3.18) verifies
K(z′, z′′, z + ξ) = K(z′ − ξ, z′′ − ξ, z).
A slightly more general form for translation invariant products of the plane, which also includes
commutative ones is represented by
(f ? g)(x) =
1
2π
∫
d2pd2qeip·xf̃(q)g̃(p− q)eα(p,q)
with f̃ , g̃ the Fourier transforms of f , g.
The function α is further restricted by the associativity request. A full analysis of the family
of translation-invariant products, together with a study of the cohomology associated to them, is
performed in [28, 81]. The usual pointwise product is reproduced by α = 0, the Moyal product
by αM(p, q) = −i/2θijqipj and the Wick–Voros product by αW(p, q) = −θq−(p+ − q+), with
q± = q1±iq2√
2
.
10 F. Lizzi and P. Vitale
4 Matrix bases for s-ordered products
In this section we consider matrix bases for the s-ordered star-products described in the previous
sections. We shall give a unified derivation for all of them and then specialize to the known
cases of the Moyal [33, 79] and Wick–Voros [49] matrix bases.
To a function on the plane R2 we associate via the quantization map (2.2) and the s-ordered
quantizer (3.14) the s-ordered operator
φ(z, z̄)→: φ̂(a, a†) :s
with : :s denoting s-ordering. This may be expanded into s-ordered powers of a, a†
φ̂ =
∑
p,q
φ̃pq :
(
a†
)p
aq :s . (4.1)
On using the number basis defined in equation (3.1) we may rewrite (4.1) as
φ̂ =
∑
p,q
φspq|p〉〈q|
with φ̃pq, φ
s
kl related by a change of basis which depends explicitly on the ordering. On applying
the dequantization formula (2.1) with the s-ordered dequantizer defined by equation (3.13) we
obtain a function in the noncommutative algebra As
φ(z, z̄) =
∑
p,q
φspqf
s
pq(z, z̄) (4.2)
with
fspq(z, z̄) = tr
(
|p〉〈q|Ûs(z̄, z)
)
=
1√
p!q!θp+q
tr
((
a†
)p|0〉〈0|aqÛs(z̄, z)).
By definition (cf. (2.3)) this yields f spq(z, z̄) as a star product
fspq(z, z̄) =
1√
p!q!θp+q
z̄p ?s f00 ?s z
q (4.3)
with fs00(z̄, z) the s-ordered symbol of the operator |0〉〈0|
fs00(z̄, z) = tr
(
|0〉〈0|Ûs(z̄, z)
)
and z̄p, zq respectively symbols of
(
a†
)p
, aq in all schemes (equation (4.3) is thus true by
definition of star product (2.3), and the use of associativity:
(fA ?s fB ?s fC)(z̄, z) = tr
(
ÂB̂ĈÛs(z̄, z)
)
with the identification  =
(
a†
)p
, B̂ = |0〉〈0|, Ĉ = aq).
It is immediate to verify that fs00(z̄, z) is idempotent independently from the particular form
of the operator Ûs. We have indeed
(f00 ?s f00)(z̄, z) = tr
(
|0〉〈0|0〉〈0|Ûs(z̄, z)
)
= f00(z̄, z).
The basis elements fpq(z, z̄) may be seen to obey the following fusion rule
fpq ?s fkl = δqkfpl (4.4)
Matrix Bases for Star Products: a Review 11
by observing that, by definition
fpq ?s fkl = tr
(
|p〉〈q|k〉〈l|Ûs(z̄, z)
)
.
This implies that every s-ordered star product may be described as matrix product. We have
indeed
φ ?s ψ(z̄, z) =
∑
φnmψpqfnm ?s fpq(z̄, z) =
∑
φnmψmqfnq(z̄, z) =
∑
(Φ ·Ψ)nqfnq(z̄, z)
with Φ, Ψ the infinite matrices with entries the series expansion coefficients of the functions φ, ψ
(see equation (4.2)). The idempotent function f00 may be computed explicitly. On using the
number basis n and the s-ordered dequantizer (3.13) we have
(1 + s)
2
f00(z̄, z) =
∑
n
〈n|0〉〈0|D̂(z)
(
s− 1
s+ 1
)a†a
D̂(−z)|n〉
= 〈0|D̂(z)
(
s− 1
s+ 1
)a†a
D̂(−z)|0〉 =
∑
nm
〈−z|n〉〈n|
(
s− 1
s+ 1
)a†a
|m〉〈m|| − z〉
= e−
z̄z
θ
∑
nm
(−1)n+m z̄nzm√
n!m!θn+m
(
s− 1
s+ 1
)m
δnm = e−
z̄z
θ e
s−1
s+1
z̄z
θ . (4.5)
As we can see, it depends explicitly on the value of s. This is a particular case of a more general
formula [16, equations (6.35), (6.36)]. We shall see in next sections that it reproduces correctly
previous results which have been obtained in specific quantization-dequantization schemes.
We can establish the useful result for the integral of the basis functions fpq. We have∫
d2zfpq(z̄, z) =
∫
d2z〈q|Us|p〉 = 2πθδpq. (4.6)
The generalization to four dimensions is readily obtained on introducing
fPQ(z̄a, za) = fp1q1(z̄1, z1) · fq2p2(z̄2, z2), a = 1, 2.
Together with the fusion rule equation (4.4), equation (4.6) ensures that the matrix basis fpq(z̄, z)
is orthogonal. This has the important consequence that the action of every field theory model
with s-ordered star product becomes a matrix action, with integrals replaced by traces. To
illustrate this point, let us consider for simplicity a scalar action with polynomial interaction, in
two dimensions
S[φ] =
∫
d2z(φ ?s Ôφ)(z̄, z) + λφ?sn(z, z̄)
with φ?sn = φ ?s φ ?s · · · ?s φ n times. On expanding the fields in the matrix basis as in in
equation (4.2) we first observe that, thanks to (4.4) repeatedly applied, the star product in the
algebra becomes an infinite-matrix product
φ1 ?s φ2 ?s · · · ?s φn(z̄, z) =
∑
pi,qi
φ1p1q1φ2p2q2 · · ·φnpnqnfp1q1 ?s fp2q2 ?s · · · ?s fpnqn
=
∑
p1,qn
(Φ1 · Φ2 · · ·Φn)p1qnfp1qn ,
where Φ = {φpq} are the infinite matrices of fields coefficients in the matrix basis expan-
sion (4.2). On integrating the latter expression by means of (4.6) we finally get∫
d2zφ1 ? φ2 ? · · · ? φn(z̄, z) = 2πθ tr(Φ1 · Φ2 · · ·Φn).
12 F. Lizzi and P. Vitale
For the kinetic term we proceed analogously, although it is in general not diagonal in the matrix
basis. Repeating the same steps we arrive at∫
d2z(φ ?s Ôφ)(z̄, z) = tr ΦOΦ
with
(O)pq;rs =
∫
d2zfpq ?s Ôfrs
the representation of the kinetic operator on the matrix basis, to be computed case by case.
Applications of this procedure may be found in Section 7.
4.1 The Moyal matrix basis
We have seen in previous sections that the Moyal product is introduced through a symmetric-
ordered quantization scheme. When qualifying R2 as the phase space of 1-dimensional systems,
the basis functions fpq(z, z̄) in this quantization scheme correspond exactly to the Wigner func-
tions associated to the density operator of the quantum oscillator states.
The Moyal matrix basis has been established long ago by J.M. Gracia-Bond́ıa and J.C. Várilly
following a slightly different approach [33, 79] with respect to the one described in previous
section. The idempotent function f00(z̄, z) has been shown [33, 79] to be the Gaussian
fM
00(z̄, z) = 2 exp(−2z̄z/θ),
which agrees with our result (4.5) at s = 0. The expression of the matrix elements φkl in terms
of φ̃pq has been computed for the Moyal case in [52].
The extension to R4
θ is straightforward. We have
φ(za, z̄a) =
∑
PQ
φPQf
M
PQ(za, z̄a) (4.7)
with a = 1, 2, P = (p1, p2) and
fM
PQ(za, z̄a) = fM
p1,q1(z1, z̄1) · fM
p2,q2(z2, z̄2).
In order to describe elements of R2
θ (resp. R4
θ), the sequences {φpq} (resp. {φ~p~q}) have to be of
rapid decay [33, 79].
4.2 The Wick–Voros matrix basis
We have seen previously that the Wick–Voros product is introduced through a weighted quan-
tization map which, in two dimensions, associates to functions on the complex plane normal
ordered operators. The inverse map which is the analogue of the Wigner map is represented by
φ(z, z̄) = 〈z|φ̂|z〉. (4.8)
The Wick–Voros product, φ ?W ψ, is particularly simple with respect to the other s-ordered
products (including the well studied Moyal one). It is defined as the expectation value over
coherent states of the operator product φ̂ψ̂. Then, for analytic functions, a very convenient way
to reformulate the quantization map (2.2) is to consider the analytic expansion
φ(z̄, z) =
∑
pq
φ̃pq z̄
pzq, p, q ∈ N,
Matrix Bases for Star Products: a Review 13
with φ̃pq ∈ C. The Wick–Voros quantizer (3.22) will then produce the normal ordered operator
φ̂ =
∑
pq
φ̃pq
(
a†
)p
aq.
We will therefore assume analyticity in what follows. The idempotent function f00 is a Gaussian,
as for the Moyal case, although with a slightly different shape
fW
00 (z̄, z) = exp(−z̄z/θ).
This result agrees with the general result (4.5) for s = 1. The basis functions fW
pq acquire the
simple form
fW
pq (z, z̄) =
e−
z̄z
θ√
p!q!θp+q
z̄pzq,
where we notice that no star product is present anymore differently from what happens in all
other situations described by equation (4.3) with s 6= 1, including the Moyal case, s = 0. This
is due to the fact that z̄p ?W f = z̄p · f as well as f ?W zq = f · zq. The generalization to R4 is
straightforward and follows the same lines as for the Moyal case. We have
fW
PQ(za, z̄a) = fW
p1,q1(z1, z̄1) · fW
p2,q2(z2, z̄2). (4.9)
5 Star products as reductions
The class of products which we have considered up to now is translation invariant, with non-
commutative parameters being constant. It is interesting to notice that, when considered in
four dimensions, through a reduction procedure these products give rise to a whole family of
star products in three dimensions, with linear noncommutativity in space coordinates. This
result was first achieved [32] by considering reductions of the Moyal product, while in [39] a par-
ticular rotation-invariant star product in three dimensions was obtained as a reduction of the
Wick–Voros product. It turns out that a reduction in terms of the Wick–Voros product is tech-
nically easier to perform, although being conceptually equivalent. We will therefore present the
reduction in such form.
The crucial step to obtain star products on F(R3), hence to deform F(R3) into a noncommu-
tative algebra, is to identify R3 with the dual, g∗, of some chosen three-dimensional Lie algebra g.
This identification induces on F(R3) the Kirillov–Poisson bracket, which, for coordinate func-
tions reads
{xi, xj} = ckijxk + bij (5.1)
with i = 1, . . . , 3 and ckij , bij , the structure constants of g. On the other hand, all three-
dimensional (Poisson) Lie algebras may be realized as subalgebras of the inhomogeneous sym-
plectic algebra isp(4), which is classically realized as the Poisson algebra of quadratic-linear
functions on R4 (C2 with our choices) with canonical Poisson bracket{
za, z̄b
}
= i, a, b = 1, 2.
It is then possible to find quadratic-linear functions
xi = xi(z
a, z̄a),
which obey (5.1). This is nothing but the classical counterpart of the Jordan–Schwinger map
realization of Lie algebra generators in terms of creation and annihilation operators [58]. Then
14 F. Lizzi and P. Vitale
one can show [32] that these Poisson subalgebras are also Wick–Voros (and Moyal) subalgebras,
that is
xi(z
a, z̄a) ?W xj(z
a, z̄a)− xj(za, z̄a) ?W xi(z
a, z̄a) = λ
(
ckijxk(z
a, z̄a) + bij
)
, (5.2)
where the noncommutative parameter λ depends on θ and shall be adjusted according to the
physical dimension of the coordinate functions xi. Occasionally we shall indicate with R3
λ the
noncommutative algebra (F(R3), ?). Equation (5.2) induces a star product on polynomial func-
tions on R3 generated by the coordinate functions xi, which may be expressed in closed form
in terms of differential operators on R3. For details we refer to [32] where all products are
classified. Here we will consider quadratic realizations of the kind
π∗(xµ) = κz̄aeabµ z
b, µ = 0, . . . , 3, (5.3)
with ei = 1
2σi, i = 1, . . . , 3 are the SU(2) generators and σi are the Pauli matrices, while e0 = 1
21.
Here we have explicitly indicated the pull-back map π∗ : F(R3) 7→ F(R4). We will shall omit
it in the following, unless necessary. κ is some possibly dimensional constant such that λ = κθ.
Notice that
x2
0 =
∑
i
x2
i .
It is possible to show that the Wick–Voros product on R4 determines the following star product
for the algebra of functions on R3, once the SU(2) generators have been chosen [39]
(φ ? ψ)(x) = exp
[
λ
2
(
δijx0 + iεkijxk
) ∂
∂ui
∂
∂vj
]
φ(u)ψ(v)|u=v=x. (5.4)
This star product implies for coordinate functions
xi ? xj = xi · xj +
λ
2
(
x0δij + iεkijxk
)
, x0 ? xi = xi ? x0 = x0xi +
λ
2
xi,
x0 ? x0 = x0
(
x0 +
λ
2
)
=
∑
i
xi ? xi − λx0,
from which one obtains
[xi, xj ]? = iλεkijxk.
Let us notice that x0 star-commutes with all elements of the algebra, so that it is possible to
define R3
λ as the star-commutant of x0.
It is possible to reduce the noncommutative algebra on R4
θ on using different three-dimensional
Lie algebras in equation (5.3) or realizations which are not even polynomial [32, 58]. These
will give different star products on R3 which are in general non-equivalent. In the following
we will just consider the star product (5.4) and refer to the corresponding noncommutative
algebra as R3
λ. The expression (5.4) for the star product in R3
λ is practically difficult to use in
calculations, for example in QFT. In next section we shall review a matrix basis for R3
λ which
makes it much easier to compute the ?-product as it will reduce the ? product (5.4) to matrix
multiplication.
Matrix Bases for Star Products: a Review 15
6 Matrix basis for R3
λ
We review a matrix basis of R3
λ which is based on a suitable reduction of the matrix basis fPQ
discussed in the previous section.
It is well known in the Jordan–Schwinger realization of the SU(2) generators, that the eigen-
values of the number operators N̂1 = a†1a1, N̂2 = a†2a2, say p1, p2, are related to the eigenvalues
of X̂2, X̂3, respectively j(j + 1) and m, by
p1 + p2 = 2j, p1 − p2 = 2m
with pi ∈ N, j ∈ N/2, −j ≤ m ≤ j, so to have
|p1p2〉 = |j +m, j −m〉 =
(
a†1
)j+m
(a2)j−m√
(j +m)!(j −m)!
|00〉,
where X̂i, i = 1, . . . , 3 are selfadjoint operators representing the su(2) Lie algebra generators on
the Hilbert space spanned by |j,m〉. Then we may relabel the matrix basis of R4
θ, equation (4.9)
as f jj̃mm̃, so to have
φ(za, z̄a) =
∑
jj̃∈N/2
j∑
m=−j
j̃∑
m̃=−j̃
φjj̃mm̃f
jj̃
mm̃(za, z̄a).
We further observe that, for φ to be in the subalgebra R3
λ we must impose j = j̃. To this it
suffices to compute
x0 ? f
jj̃
mm̃ − f
jj̃
mm̃ ? x0 = λ(j − j̃)f jj̃mm̃
and remember that R3
λ may be alternatively defined as the ?-commutant of x0. This requires
j = j̃.
We have then
φ(xi) =
∑
j
j∑
m,m̃=−j
φjmm̃v
j
mm̃
with
vjmm̃ := f jjmm̃ =
z̄j+m1 f00(z̄1, z1)zj+m̃1 z̄j−m2 f00(z̄2, z2)zj−m̃2√
(j +m)!(j −m)!(j + m̃)!(j − m̃)!θ4j
.
The orthogonality property now reads
vjmm̃ ? vj̃nñ = δjj̃δm̃nv
j
mñ.
As for the normalization we have∫
d2z1d2z2v
j
mm̃(z, z̄) = 4π2θ2δmm̃.
The star product in R3
λ becomes a matrix product
φ ? ψ(x) =
∑
φj1m1m̃1
φj2m2m̃2
vj1m1m̃1
? vj2m2m̃2
=
∑
φj1m1m̃1
φj2m2m̃2
vj1m1m̃2
δj
1j2δm̃1m2
16 F. Lizzi and P. Vitale
while the integral may be defined through the pullback to R4
θ∫
R3
λ
d3xφ ? ψ := κ2
∫
R4
θ
d4xπ?(φ) ?M π∗(ψ) = 4π2λ2 tr ΦΨ
hence becoming a trace.
In analogy with the present derivation, the matrix basis adapted to the Moyal product [33, 79]
has been reduced to three dimensions in [69] where applications to quantum mechanics (the
hydrogen atom) are considered.
7 Field theories on noncommutative spaces as matrix models
Field theories on noncommutative spaces based on the Moyal product were introduced in [61, 72].
It was soon realized that they could be very effectively described by matrix models. For example
it was shown in [3] that defining a noncommutative field theory on a noncommutative torus
(which we discuss below in Section 8.1), the theory is defined on a lattice and becomes a matrix
model of the IKKT type [42]. Another application of matrix bases shows that the gauge Lie
algebra of a theory defined with the Moyal product is a particular form of SU(∞) [52] related
to the inner automorphisms of the underlying deformed algebra of functions on spacetime.
In several relevant examples of QFT on noncommutative spaces the introduction of an or-
thogonal matrix basis has made it possible to explicitly compute the propagator and the vertices
of the models investigated. We describe some of them in this section. The importance of the
matrix basis is that a perturbative analysis becomes possible, reducing the problem of loop cal-
culations to taking traces, which, once regularized with the introduction of a cutoff (so to have
finite matrices). These may be implemented with a computer program, and some steps in this
direction have been taken in [74].
7.1 The Grosse–Wulkenhaar model
An important application of the matrix basis for the Moyal plane is the perturbative analysis of
the Grosse–Wulkenhaar harmonic model [36, 37]. For simplicity we shall only review here the
two-dimensional case [36] to illustrate the procedure.
The model deals with a scalar theory with quartic interaction. It is described by the action
S =
∫
d2z
(
∂zφ ? ∂z̄φ+ 2
Ω2
θ2
(zφ ?M z̄φ+ z̄φ ?M zφ) +
1
2
µ2
0φ ?M φ+
λ
4!
φ?M4
)
. (7.1)
The harmonic term is crucial in four dimensions to cure the famous UV/IR mixing [61], which
is quadratic in d = 4. It is however worth noting that it breaks the translation invariance of the
action.
On using the series expansion (4.7) for the fields and the orthogonality properties of the
matrix basis described in Section 4.1, together with equations (3.26), we may rewrite the ac-
tion (7.1) as
S = Skin + Sint
with
Sint =
λ
4!
πθ tr Φ · Φ · Φ · Φ, Skin = tr Φ∆Φ. (7.2)
Matrix Bases for Star Products: a Review 17
Hence we observe that, while the interaction term is polynomial in the matrix Φ ≡ (φmn), the
kinetic term is highly non-local (non-diagonal), with
∆mn,kl =
(
µ2
0 + 2
(1 + Ω2)
θ
(m+ n+ 1)
)
δnkδml − 2
(1− Ω2)
θ
√
(n+ 1)(m+ 1)δn+1,kδm+1,l
− 2
(1− Ω2)
θ
√
nmδn−1,kδm−1,l. (7.3)
The propagator denoted by Pmn;kl is the inverse of the kinetic term. It is defined by∑
k,l
∆mn;klPlk;sr = δmrδns,
∑
k,l
Pnm;lk∆kl;rs = δmrδns. (7.4)
∆ satisfies an index conservation law
∆mn;kl 6= 0 ⇐⇒ m+ n = k + l.
This implies that equation (7.3) depends only on three indices. Therefore, setting n = α −m,
k = α− l, with α = m+ n = k + l we set
∆m,α−m;α−l,l := ∆
(α)
m,l. (7.5)
One observes that, for each value of α, ∆
(α)
ml is an infinite real symmetric tridiagonal matrix
which can be related to a Jacobi operator. Therefore, the diagonalization of (7.5) can be
achieved by using a suitable family of Jacobi orthogonal polynomials. This is a general feature
of all subsequent models which shall be analyzed in this section.
Denoting generically by λk, k ∈ N the eigenvalues of ∆
(α)
mn (7.5), we write it as
∆
(α)
ml =
∑
p∈N
R(α)
mp
(
2(1 + Ω2)
θ
λp + µ2
0
)
R(α)†
pl
with ∑
p∈N
R(α)
mpR(α)†
pl =
∑
p∈N
R(α)†
mpR(α)
pl = δml, (7.6)
where R(α)†
mn = R(α)
nm. Then, combining with equation (7.3) we obtain the following 3-term
recurrence relation(
1− Ω2
)√
(m+ 1)(α+m+ 1)R(α)
m+1(λ) +
(
1− Ω2
)√
m(α+m)R(α)
m−1(λ)
+
(
λ−
(
1 + Ω2
)
(α+ 1 + 2m)
)
R(α)
m (λ) = 0, ∀m, q ∈ N,
where we have traded the discrete index q for λ. On introducing a cutoff N on the matrix
indices, it has been shown in [36] that this is the recurrence equation for modified Laguerre
polynomials [45] Lα,ωm (λ) with ω1/2 = (1 − Ω2)/(1 + Ω2). The eigenvalues of the Laplacian
are the zeroes of the modified Laguerre polynomials. The eigenfunctions of the Laplacian
are therefore proportional to modified Laguerre polynomials, up to a normalization function,
f(N,α,m), which is determined from the orthonormality request.
Once we have diagonalized the kinetic operator, the propagator is readily obtained. From
equation (7.4) we obtain
P (N,α,ω)
mn =
N∑
p=0
f2(N,m,α)Lα,ωm (λp)
1
2 (1+Ω2)
θ λ+ µ2
0
Lα,ωn (λp).
18 F. Lizzi and P. Vitale
The limitN →∞ is easy to perform in the case ω = 1 where the product of Laguerre polynomials
gives rise to the integration measure (see [36] for details).
In [37] the whole analysis has been repeated for the Moyal space R4
θ. The kinetic term
is of the same kind as the one considered here, although in higher dimensions. It turns out
that the recurrence relation which is relevant there, is satisfied by another family of orthogonal
polynomials, the so called Meixner polynomials [45].
7.2 The translation invariant model
The translation invariant model has been introduced in [38]. Its importance resides in the fact
that it is renormalizable, while preserving translation invariance. Indeed, as already noticed in
the previous sections, the Moyal star product is an instance of a translation invariant product,
according to (3.27). This implies that every commutative translation invariant theory keeps
such an invariance when deformed by the sole replacement of the commutative star-product
with the Moyal star product or any other translation-invariant one. However we have seen in
the previous section that the noncommutative λφ4 field theory is not renormalizable, unless
a translation invariance breaking term is introduced. On the other hand, the model briefly
described below has the advantage of modifying the propagator of the λφ4 model, without
destroying the symmetries of its commutative analogue.
The action which describes the model in four dimensions (Euclidean) is
S =
∫
d4x
[
1
2
(
∂µφ ? ∂µφ+
a
θ2
∂−1
µ φ ? ∂−1
µ φ+m2φ ? φ
)]
+
λ
4!
φ?4
with
∂−1
µ φ(x) =
∫
dxµφ(x) =
∫
dp
1
ipµ
φ̃(p)eip·x
the antiderivative and φ̃(p) the Fourier transform of φ(x). In [78] the model has been studied
with a generic translation invariant product showing that the universal properties do not depend
on the particular product of the family. In the same paper the model is formulated in the Wick–
Voros matrix basis of Section 4.2. The kinetic term was computed but it was not recognized that
it is of the same kind as the Grosse–Wulkenaar one, that is, an operator of Jacobi type, while
the interaction term is the same as in (7.2). Therefore, the propagator can be found with the
same techniques as in the Grosse–Wulkenhaar model. This point deserves further investigation.
We shall come back to this issue elsewhere.
7.3 Gauge model on the Moyal plane
The UV/IR mixing also occurs in gauge models on 4-dimensional Moyal space [40, 60]. For early
studies, see e.g. [13, 14] and references therein. The mixing appears in the naive noncommutative
version of the Yang–Mills action given by S0 = 1
4
∫
d4x(Fµν ? Fµν)(x), showing up at one-loop
order as a hard IR transverse singularity in the vacuum polarization tensor. Attempts to extend
the Grosse–Wulkenhaar harmonic solution to a gauge theoretic framework have singled out
a gauge invariant action expressed as [22]
SΩ =
∫
ddx
(
1
4
Fµν ? Fµν +
Ω2
4
{Aµ,Aν}2? + κAµ ?Aµ
)
, (7.7)
where Ω and κ are real parameters, while Aµ = Aµ − Ainv
µ is a gauge covariant one-form given
by the difference of the gauge connection and the natural gauge invariant connection
Ainv
µ = −θ−1
µν x
ν .
Matrix Bases for Star Products: a Review 19
Unfortunately, the action (7.7) is hard to deal with when it is viewed as a functional of the
gauge potential Aµ. This is mainly due to its complicated vacuum structure explored in [23].
When expressed as a functional of the covariant one-form Aµ, the action (7.7) bears some
similarity with a matrix model, where the field Aµ can be represented as an infinite matrix in
the Moyal matrix base.
We will review here the two-dimensional case [59] and we shall consider fluctuations around
a particular vacuum solution, which shall make the kinetic term of the action into a Jacobi type
operator, as in the model considered in Section 7.1. This choice makes the model tractable and
permits to invert for the propagator.
We set
A =
A1 + iA2√
2
, A† =
A1 − iA2√
2
.
Then, one obtains
SΩ[A] =
∫
d2x
((
1 + Ω2
)
A ?A† ?A ?A† +
(
3Ω2 − 1
)
A ?A ?A† ?A† + 2κA ?A†
)
.
The star product used here is the Moyal star product, although any star product of the equiv-
alence class (translation invariant ones) would give the same results. This action shares some
similarities with the 6-vertex model although the entire analysis relies on the choice of a vacuum
around which we shall perform fluctuations.
The strategy used is standard: one chooses a particular vacuum (the background), expand
the action around it, fix the background symmetry of the expanded action.
From the perspective of the present review an interesting feature of this model is the fact
that, when a particular non-trivial vacuum is chosen, among those classified in [23], the kinetic
term of the action becomes a Jacobi type operator, therefore invertible for the propagator in
terms of orthogonal polynomials. We therefore refer for details to [59] and we concentrate here
on the form of the kinetic operator, when the special vacuum is chosen. In the Moyal basis the
vacuum is expressed as
Z(x) =
∑
m,n∈N
Zmnfmn(x)
with
Zmn = − i
2
√
−3κδm+1,n, κ < 0, ∀m,n ∈ N.
This latter expression is a solution of the classical equation of motion for Ω2 = 1
3 . When
expanded around this vacuum the kinetic part of the action becomes
Skin[φ] =
∑
m,n,k,l∈N
φmnφkl∆mn;kl,
where φ =
∑
mn
φmnfmn are the gauge field fluctuations expanded in the Moyal matrix basis. The
kinetic operator reads
∆
(1/3)
mn;kl = (−κ)(2δmlδnk − δk,n+1δm,l+1 − δn,k+1δl,m+1),
and satisfies ∆
(1/3)
mn;kl 6= 0 ⇐⇒ m+ n = k + l.
The propagator, Pmn;kl, is defined as in (7.4). Proceeding as in the Grosse–Wulkenhaar case,
we pose α = m+ n = k + l so that
∆
(1/3)
m,α−m;α−l,l := ∆α
m,l = µ2(2δml − δm,l+1 − δl,m+1), ∀m, l ∈ N, (7.8)
20 F. Lizzi and P. Vitale
where µ2 = −κ. Notice that in this case it does not depend on α. Therefore, we set ∆α
m,l = ∆ml
to simplify the notations.
One observes that ∆ml is an infinite real symmetric tridiagonal matrix which can be related
to a Jacobi operator. Therefore, the diagonalization of (7.8) can be achieved by using a suitable
family of Jacobi orthogonal polynomials.
We thus arrive at the following recurrence equation
Rm+1(x) +Rm−1(x) = (2 + x)Rm(x), ∀m ∈ N,
where we have posed x = −λq. On restricting to N ×N submatrices, it is possible to show that
the recurrence equation above is satisfied by Chebyschev polynomials of second kind [45]:
Um(t) := (m+ 1) 2F1
(
−m,m+ 2;
3
2
;
1− t
2
)
, ∀m ∈ N,
where 2F1 denotes the hypergeometric function. Moreover, the eigenvalues are exactly given by
the roots of RN (x). So we have
Rm(x) = f(x)Um
(
2 + x
2
)
, ∀m ∈ N,
where f(x) is a normalization function to be determined by the orthonormality condition (7.6).
The eigenvalues of ∆N
ml are now entirely determined by the roots of UN (t). These are given by
tNk = cos( (k+1)π
N+1 ), k = 0, 2, . . . , N − 1. Then, the eigenvalues for the kinetic operator ∆N
ml are
µ2λNk = 2µ2
(
1− cos
(
(k + 1)π
N + 1
))
, k ∈ {0, 2, . . . , N − 1},
and satisfy for finite N
0 < µ2λkN < 4µ2.
Thus, we have obtained:
RNmq = f(N, q)Um
(
tNq
)
= f(N, q)
sin
[π(m+1)(q+1)
N+1
]
sin
[π(q+1)
N+1
] , 0 ≤ m, q ≤ N − 1, (7.9)
where we used Um(cos θ) = sin((m+1)θ)
sin θ .
The normalization function is found to be
f(N,m) =
(
(−1)m(N + 1)
sin
[N(m+1)π
N+1
]
sin3
[ (m+1)π
N+1
] )−
1
2
, 0 ≤ p, m ≤ N − 1.
Once we have the polynomials which diagonalize the kinetic term we can invert for the propaga-
tor. Keeping in mind equations (7.4) and (7.8), we set Pmn := Pm,α−n;α−l,l where α = m+ n =
k + l. It follows from the above that for fixed N the inverse of ∆N
mn denoted by PNmn can be
written as
PNmn =
1
2µ2
N−1∑
p=0
f2(N, p)Um
(
tNp
) 1
1− tNp
Un
(
tNp
)
. (7.10)
Matrix Bases for Star Products: a Review 21
Taking the limit N →∞, the comparison of the relation δml =
∑
p
RNmpRNlp where the RNmn’s are
given by equation (7.9) to the orthogonality relation among the Chebyshev polynomials Un∫ 1
−1
dµ(x)Um(x)Un(x) =
π
2
δmn, dµ(x) = dx
√
1− x2,
permits one to trade the factor f2(N, p) in PNmn (7.10) for the compactly supported integration
measure dµ(x).
We finally obtain the following rather simple expression for the inverse of the kinetic opera-
tor (7.8)
Pmn;kl = δm+n,k+lPml, Pml =
1
πµ2
∫ 1
−1
dx
√
1 + x
1− x
Um(x)Ul(x). (7.11)
Notice that the integral in (7.11) is well-defined leading to finite Pml when m and l are finite.
7.4 The scalar model on R3
λ
In this section we review a family of scalar field theories on R3
λ, the noncommutative algebra
introduced in Section 5. The contents and presentation are based on [82], where one loop
calculations were performed. Here the stress will be, as for the models described in the previous
sections, on the use of a matrix basis (in this case the one of Section 6) to obtain a non-local
matrix model and show that the kinetic term of the theory is of Jacobi type, so that it can be
inverted for the propagator using standard techniques of orthogonal polynomials.
Let us recall that the algebra R3
λ is generated by the coordinate functions xµ, µ = 0, . . . , 3.
The coordinate x0 is in the center of the algebra and plays the role of the radius of fuzzy
two-spheres which foliate the whole algebra.
Let
S[φ] =
∫
φ ?
(
∆ + µ2
)
φ+
g
4!
φ ? φ ? φ ? φ, (7.12)
where ∆ is the Laplacian defined as
∆φ = α
∑
i
D2
i φ+
β
κ4
x0 ? x0 ? φ (7.13)
and
Di = κ−2[xi, · ]?, i = 1, . . . , 3 (7.14)
are inner derivations of R3
λ. The mass dimensions are [φ] = 1
2 , [g] = 1, [Di] = 1. α and β are
dimensionless parameters.
The second term in the Laplacian has been added in order to introduce radial dynamics.
From (5.4) we have indeed
[xi, φ]? = −iλεijkxj∂kφ
so that the first term, that is [xi, [xi, φ]?]? can only reproduce tangent dynamics on fuzzy spheres;
this is indeed the Laplacian usually introduced for quantum field theories on the fuzzy sphere
(cf. Section 8.2). Whereas
x0 ? φ = x0φ+
λ
2
xi∂iφ
contains the dilation operator in the radial direction.
22 F. Lizzi and P. Vitale
Therefore, the highest derivative term of the Laplacian defined in (7.13) can be made into the
ordinary Laplacian on R3 multiplied by x2
0, for the parameters α and β appropriately chosen.
For simplicity, we restrict the analysis to α, β positive, which is a sufficient condition for the
spectrum to be positive.
It is not difficult to verify that the following relations old
[x+, [x−, v
j
mm̃]?]? = λ2
{(
(j +m)(j −m+ 1) + (j + m̃+ 1)(j − m̃)
)
vjmm̃
−
√
(j +m)(j −m+ 1)(j + m̃)(j − m̃+ 1)vjm−1m̃−1
−
√
(j +m+ 1)(j −m)(j + m̃+ 1)(j − m̃)vjm+1m̃+1
}
,
[x−, [x+, v
j
mm̃]?]? = λ2
{(
(j +m+ 1)(j −m) + (j + m̃)(j − m̃+ 1)
)
vjmm̃
−
√
(j +m)(j −m+ 1)(j + m̃)(j − m̃+ 1)vjm−1m̃−1
−
√
(j +m+ 1)(j −m)(j + m̃+ 1)(j − m̃)vjm+1m̃+1
}
,[
x3,
[
x3, v
j
mm̃
]
?
]
?
= λ2(m− m̃)2vjmm̃,
x0 ? x0 ? v
j
mm̃ = λ2j2vjmm̃.
On expanding the fields in the matrix basisφ =
∑
j,mm̃
φjmm̃v
j
mm̃ we rewrite the action in (7.12) as
a matrix model action
S[φ] = κ3
{
tr(Φ(∆(α, β))Φ) +
g
4!
tr(ΦΦΦΦ)
}
, (7.15)
where sums are understood over all the indices and tr :=
∑
j
trj . The kinetic operator may be
computed to be
(∆(α, β))j1j2m1m̃1;m2m̃2
=
1
π2θ2
∫
vj1m1m̃1
? (∆(α, β))vj2m2m̃2
=
λ2
κ4
δj1j2
{
δm̃1m2δm1m̃2D
j2
m2m̃2
− δm̃1,m2+1δm1,m̃2+1B
j2
m2,m̃2
− δm̃1,m2−1δm1,m̃2−1H
j2
m2,m̃2
}
with
Dj
m2m̃2
=
[
(2α+ β)j2 + 2α(j2 −m2m̃2)
]
+ λ2µ2,
Bj
m2m̃2
= α
√
(j +m2 + 1)(j −m2)(j + m̃2 + 1)(j − m̃2),
Hj
m2m̃2
= α
√
(j +m2)(j −m2 + 1)(j + m̃2)(j − m̃2 + 1).
Let us notice that the use of the matrix basis vjmn yields an interaction term which is diagonal
whereas the kinetic term is not diagonal. Had we used the expansion of φ in the fuzzy harmonics
base (Y j
lk), j ∈
N
2 , l ∈ N, 0 ≤ l ≤ 2j, −l ≤ k ≤ l (see Section 8.2), we would have obtained
a diagonal kinetic term with a non-diagonal interaction term. The latter will be the choice in
Section 8.2 where we follow the traditional approach to the study of the fuzzy sphere Laplacian.
Moreover, we observe that the action (7.15) is expressed as an infinite sum of contributions,
namely S[Φ] =
∑
j∈N
2
S(j)[Φ], where the expression for S(j) can be read off from (7.15) and describes
a scalar action on the fuzzy sphere Sj .
We now pass to the calculation of the propagator, through the inversion of the kinetic term
in the action. Because of the remark above, this is expressible into a block diagonal form.
Explicitly
SKin[Φ] = κ3
∑
j
∑
m,m̃
φj1m1m̃1
(∆)j1j2m1m̃1;m2m̃2
φj2m2m̃2
.
Matrix Bases for Star Products: a Review 23
Since the mass term is diagonal, let us put it to zero for the moment. We shall restore it at the
end. One has the following law of indices conservation
∆j1j2
mn;kl 6= 0 =⇒ j1 = j2, m+ k = n+ l.
The inverse of ∆j1j2
mn;kl(α, β) is thus defined by
j2∑
k,l=−j2
∆j1j2
mn;lkP
j2j3
lk;rs = δj1j3δmsδnr,
j2∑
m,n=−j2
P j1j2rs;mn∆j2j3
mn;kl = δj1j3δrlδsk,
for which the law of indices conservation still holds true as
P j1j2mn;kl 6= 0 =⇒ j1 = j2, m+ k = n+ l.
To determine P j1j2mn;kl one has to diagonalize ∆j1j2
mn;kl along the same lines as in previous sections, by
means of orthogonal polynomials. This is done in detail in [82] where the orthogonal polynomials
are found to be the dual Hahn polynomials. Here however we take a shortcut, because we already
know an alternative orthogonal basis for R3
λ where the kinetic part of the action is diagonal, that
is the fuzzy spherical harmonics. It can be shown that dual Hahn polynomials and the fuzzy
spherical harmonics are indeed the same object, up to a proportionality factor.
7.4.1 The kinetic action in the fuzzy spherical harmonics base
Fuzzy Spherical Harmonics Operators, are, up to normalization factors, irreducible tensor ope-
rators
Ŷ j
lk ∈ End(Vj), l ∈ N, 0 ≤ l ≤ 2j, −l ≤ k ≤ l,
whereas the unhatted objects Y j
lk are their symbols and are sometimes referred to as fuzzy
spherical harmonics with no other specification (notice however that the functional form of
the symbols does depend on the dequantization map that has been chosen). Concerning the
definition and normalization of the fuzzy spherical harmonics operators, we use the following
conventions. We set
J± =
x̂±
λ
.
We have, for l = m,
Ŷ j
ll := (−1)l
√
2j + 1
l!
√
(2l + 1)!(2j − l)!
(2j + l + 1)!
(J+)l
while the others are defined recursively through the action of J−
Ŷ j
lk := [(l + k + 1)(l − k)]−
1
2 [J−, Ŷ
j
l,k+1],
and satisfy(
Ŷ j
lk
)†
= (−1)k−2j Ŷ j
l,−k, 〈Ŷ j
l1k1
, Ŷ j
l2k2
〉 = tr
((
Ŷ j
l1k1
)†
Ŷ j
l2k2
)
= (2j + 1)δl1l2δk1k2 .
The symbols are defined through the dequantization map (4.8)
Y j
lk := 〈z| Ŷ j
lk |z〉. (7.16)
24 F. Lizzi and P. Vitale
We have then
[xi, [xi, Y
j
lk]?]? = λ2〈z|[Ji, [Ji, Ŷ j
lk]]|z〉 = λ2l(l + 1)Y j
lk.
In order to evaluate the action of the full Laplacian (7.13) on the fuzzy spherical harmonics we
need to compute x0 ? Y
j
lk. To this we express the fuzzy spherical harmonics in the canonical
base vjmm̃
Y j
lk =
∑
−j≤m,m̃≤j
(
Y j
lk
)
mm̃
vjmm̃,
where the coefficients are given in terms of Clebsch–Gordan coefficients by
(Y j
lk)mm̃ = 〈v̂jmm̃|Ŷ
j
lk〉 =
√
2j + 1(−1)j−m̃
(
j j l
m −m̃ k
)
, −j ≤ m, m̃ ≤ j,(
Y j
lk
†)
mm̃
= (−1)−2j
(
Y j
lk
)
m̃m
.
We have then
x0 ? Y
j
lk =
∑
−j≤m,m̃≤j
(
Y j
lk
)
mm̃
x0 ? v
j
mm̃ = λjY j
lk.
Thus we verify that in the fuzzy spherical harmonics base the whole kinetic term is diagonal,
∆(α, β)Y j
lk =
λ2
κ4
(
αl(l + 1) + βj2
)
Y j
lk, j ∈ N
2
, 0 ≤ l ≤ 2j, l ∈ N, −l ≤ k ≤ l,
with eigenvalues
λ2
κ4
γ(j, l;α, β) :=
λ2
κ4
(
αl(l + 1) + βj2
)
.
We can expand the fields φ ∈ R3
λ in the fuzzy harmonics base φ =
∑
j∈N
2
2j∑
l=0
l∑
k=−l
ϕjlkY
j
lk, with the
coefficients ϕjlk related to those in the canonical base φjmm̃ by a change of basis.
Therefore, we can compute the kinetic action in the fuzzy harmonics base to be∫
φ ? (∆ + µ2)φ =
λ2
κ
∑
|ϕjlk|
2(2j + 1)
(
γ(j, l;α, β) +
κ4
λ2
µ2
)
,
which is positive for α, β ≥ 0. We define for further convenience
(∆diag)j1j2l1k1l2k2
=
1
λ3
∫
Y j1
l1k1
?∆(α, β)Y j2
l2k2
=
1
λ2
(−1)k1+2j1(2j1 + 1)γ(j1, l1;α, β)δj1j2δl1l2δ−k1k2 .
Then, the kinetic term in the canonical basis may be expressed in terms of the diagonal one
∆j1j2
m1m̃1m2m̃2
=
1
(2j1 + 1)2
(
Y j1
l1k1
)
m1m̃1
(
∆j1j2
diag
)
l1k1l2k2
(
Y j2
l2k2
)
m2m̃2
.
The propagator is then
[P j1j2 ]m1m̃1m2m̃2 =
(
Y j1
l1k1
)
m1m̃1
[(
∆j1j2
diag
)−1]
l1k1l2k2
(
Y j2
l2k2
)
m2m̃2
.
Matrix Bases for Star Products: a Review 25
On replacing the expression for the diagonal inverse we finally obtain
[P j1j2 ]m1m̃1m2m̃2 = (−1)−k+2j1δj1j2
×
2j1∑
l=0
l∑
k=−l
κ4
λ2
1
(2j1+ 1)
(
γ(j1, l;α, β) + µ2
)(Y j1†
lk
)
m1m̃1
(
Y j2
lk
)
m2m̃2
. (7.17)
In [82] one loop calculations have been performed showing the absence of divergences.
We finally that these results may be generalized to gauge theories on R3
λ. In [30] the following
gauge model has been considered
Scl(Ai) = tr
(
αAiAjAjAi + βAiAjAiAj + ΘεijkAiAjAk +mAiAi
)
, (7.18)
where Ai = −iAi + ηi and α, β, Θ, m are real parameters. Ai is the gauge potential and ηi is
the invariant connection associated to the differential calculus on R3
λ
η(Di) := ηi =
i
κ2
xi,
where Di are the inner derivations of the algebra introduced in equation (7.14). By requiring
that no linear terms in Ai be involved, the action (7.18) may be rewritten as
Scl(Ai) = tr
(
F †ijFij + γ(εijkAiAjAk +
3
2
λ
κ2
AiAi)
)
.
with appropriately defined parameters. The total action is thus rewritten as the sum of a Yang–
Mills and a Chern–Simons term,
Scl(Ai) = trF †ijFij + SCS
cl (Ai),
which is of the same form as the Alekseev–Recknagel–Schomerus gauge action on the fuzzy
sphere [2], although here we have a sum over all fuzzy spheres of the foliation of R3
λ.
The model has been studied in the matrix basis of R3
λ showing that, when the action is
formally massless, the gauge and ghost propagators are of the same form as the scalar propa-
gator (7.17) found above. This result has been used to perform one loop calculations. It is
found that the infrared singularity of the propagator stemming from masslessness disappears
from the computation of the correlation functions. Moreover it is shown that this massless gauge
invariant model on R3
λ has quantum instabilities of the vacuum, signaled by the occurrence of
non vanishing tadpole (1-point) functions for some but not all of the components of the gauge
potential.
We close this section observing that all the models considered are connected to Jacobi type
kinetic operators, which give rise to three term recurrence equations. These are solved for
specific families of polynomials which allow in turn to determine the propagator. A systematic
analysis is performed in [31]. However, it is interesting to notice that five terms recurrence
relations emerge, for example in the two-dimensional gauge model, with a different choice for
the vacuum, but also in the three-dimensional gauge model on R3
λ, in the massive case, which
might be worth to investigate.
8 Fuzzy spaces
Fuzzy spaces are matrix approximations of ordinary spaces. Their importance lies in the fact
that, although the algebra which approximates the original functions on the space is finite-
dimensional, the original group of isometries is preserved. The literature on fuzzy spaces is vast
(see [6, 54] and references therein), in this section we will limit ourselves to a presentation of
these fuzzy space which shows how they can be interpreted in a way similar to the matrix basis
for the ? products in the previous section.
26 F. Lizzi and P. Vitale
8.1 The fuzzy torus
The fuzzy torus is a finite-dimensional of the noncommutative torus [67], which is probably
the most studied noncommutative space. It is in some sense a compact version of the Moyal
plane introduced in the previous section. Consider the algebra of functions on a two-dimensional
torus1. In Fourier transform they can be represented as
f =
∞∑
n1,n2=−∞
fn1n2e2πin1x1e2πin2x2 , (8.1)
where we impose that the fn1n2 decrease exponentially as ni → ±∞. The noncommutative torus
is obtained with the substitution e2πixi → Ui with the condition
U1U2 = e2πiθU2U1. (8.2)
Loosely speaking this is what would be obtained imposing the commutation relation (3.10),
except that of course the x’s are not well defined quantities on a torus. One can represent the
Ui’s as operators on the Hilbert space of L2(S1) functions a circle as follows:
U1ψ(α) = ψ(α+ 2πθ), U2ψ(α) = e2παψ(α).
It is immediate to verify relation (8.2). It follows that the noncommutative torus is given by
a Weyl map
Ω̂(f) =
∞∑
n1,n2=−∞
fn1n2U
n1
1 Un2
2
giving rise to the noncommutative ? product defined by the twisted convolution of Fourier
coefficients:
(f ∗ g)n1n2 =
∞∑
m1,m2=−∞
fm1m2gn1−m1,n2−m2e2πi(n1m2−n2m1).
The representation of the operators U1 and U2 in the discrete basis of L2(S1) given by ϕp = e2πipα
is given by
U1pq = δp,q−1, U2pq = e2πpδp,q (no sum over p).
In the rational case of θ = M
N it is possible to find a finite (N×N)-dimensional representation
of the Ui’s as:
U
(N)
1 =
1
e2πiθ
e2πiθ
. . .
e2πi(N−1)θ
, U
(N)
2 =
0 1 0
0 1
. . .
. . .
. . . 1
1 0
.
They are unitary and traceless (since
N−1∑
k=0
e2πiθk = 0), satisfy
UMi = IN
1Higher-dimensional cases can be studied, but in this review we will confine ourselves to two dimensions.
Matrix Bases for Star Products: a Review 27
and obey the commutation relation (8.2). The algebra generated by the U
(N)
i with the expan-
sion (8.1) is called a fuzzy torus. Note that a noncommutative torus with rational θ = M
N is not
the same algebra of a fuzzy torus with the same θ. The former is infinite-dimensional, while
the latter if finite-dimensional. The fuzzy torus can however be seen an an approximation of
a noncommutative torus, by taking the value of N , and hence the size of the matrices larger and
larger. Since any irrational number can be approximated arbitrarily by a sequence of rationals,
for example using continuous fractions, there is a sequence of finite-dimensional algebras which
approximate the algebra of the noncommutative torus. The appropriate tool for this approxima-
tion is the inductive limit [15], but the infinite-dimensional algebra of the noncommutative (or
commutative) torus is not approximable by a sequence of finite-dimensional algebras. It is how-
ever possible [46, 65] to prove that the inductive limit of a sequence of these finite-dimensional
algebras converges to a larger algebra which contains the noncommutative tours, as well as the
algebra of all the tori which are Morita equivalent to it.
8.2 The fuzzy sphere
The fuzzy sphere [41, 53] is the most famous example of fuzzy space, and is usually presented
using the identification xi ∝ Ji, where the x’s are the coordinates on R3 and the J ’s the
generators of angular momentum in a particular representation. The sphere constraint
∑
i
x2
i =
R2 is then equivalent to the Casimir relation
∑
i
J2
i ∝ I2j+1. In this section we will present the
fuzzy sphere as an example of Weyl–Wigner correspondence and an instance of a star product.
The product is based, as in the case of the Wick–Voros plane, on the use of coherent states. Since
the sphere is a coadjoint orbit of SU(2) the relevant coherent states are the generalization of the
usual ones pertaining to the one related this group. Notice however, that it could be equivalently
considered as a noncommutative subalgebra of R3
λ (cf. Section 6 at some fixed value of x0).
Consider SU(2) in a particular representation. The construction can be made for every Lie
algebra [64], and related coadjoint orbits. Consider a representation of the group on the finite-
dimensional Hilbert space H2j+1:
g ∈ SU(2)→ U(g),
where U(g) is a 2j+ 1×2j+ 1 matrix, j ∈ Z/2. Consider a vector |ψ〉. A subgroup Hψ ⊂ SU(2)
will leave it invariant up to a phase. Consider now a fiducial vector |ψ0〉 such that Hψ0 is
maximal. A natural choice for the fiducial state is a highest weight vector for the representation
|ψ0〉 = |j, j〉, where we use the basis |j,m〉 of simultaneous eigenvectors of J2 and J3, with
m = −j,−j + 1, . . . , j. The sphere is the quotient of SU(2), which topologically is a three
sphere S3, by the subgroup Hψ0 , which in this case is U(1),
S2 = SU(2)/Hψ0 .
Consider the usual basis of S3 given by the three Euler angles α ∈ [0, 4π), β ∈ [0, π),
γ ∈ [0, 2π). the corresponding element in SU(2) is given by
U(α, β, γ) = e−iαJ3e−iβJ2e−iγJ3 .
The points for which β = 0 are left invariant up to a phase. The sphere S2 can therefore
be characterized by the coordinates β and α mod 2π, which we may identify with the usual
coordinates on the sphere ϑ and ϕ.
Choosing a representative element g in each equivalence class of the quotient, the set of
coherent states is defined by
|ϑ, ϕ〉2j+1 = U(g)|j, j〉.
28 F. Lizzi and P. Vitale
They depend on the dimension of the representation. Projecting onto the basis elements |j,m〉
one finds
|ϑ, ϕ〉2j+1 =
j∑
m=−j
√
(2j)!
(j +m)!(j −m)!
cosj+m
ϑ
2
sinj−m
ϑ
2
e−imϕ |j,m〉,
As in the earlier case, coherent states are non-orthogonal and overcomplete
2j+1〈ϑ′, ϕ′|ϑ, ϕ〉2j+1 = e−ij(ϕ′−ϕ)
[
ei(ϕ′−ϕ) cos
ϑ
2
cos
ϑ′
2
+ sin
ϑ
2
sin
ϑ′
2
]2j
,
I =
2j + 1
4π
∫
S2
dΩ|ϑ, ϕ〉2j+1 2j+1〈ϑ, ϕ|,
where dΩ = sinϑdϑdϕ.
As in the case discussed in Section (3.3) we can use coherent states to define a map from
operators to functions. Note that in this case, as in the fuzzy torus case, the map is not one-to-
one.
F̂ (2j+1) ∈M2j+1(C) 7−→ f (2j+1) ∈ C
(
S2
)
,
f (2j+1)(ϑ, ϕ) = 2j+1〈ϑ, ϕ|F̂ (2j+1)|ϑ, ϕ〉2j+1. (8.3)
This is also called the Berezin symbol of the matrix [12].
Spherical harmonics operators, already introduced in Section 7.4.1, form a basis for the
algebra of 2j + 1×2j + 1 matrices. Therefore elements F̂ (2j+1) ∈M2j+1(C) can be expanded as
F̂ (2j+1) =
j∑
l=0
l∑
m=−l
F
(2j+1)
lm Ŷ
(2j+1)
lm ,
with coefficients
F
(2j+1)
lm =
tr
[
Ŷ
(2j+1)†
lm F̂ (2j+1)
]
tr Ŷ
(2j+1)†
lm Ŷ
(2j+1)
lm
.
Fuzzy harmonics are defined as the symbols of spherical harmonics operators over coherent
states (cf. our previous definition (7.16))
2j+1〈ϑ, ϕ|Ŷ (2j+1)
lm |ϑ, ϕ〉2j+1 = Y
(2j+1)
lm (ϑ, ϕ). (8.4)
They form a basis in the noncommutative algebra C(S2). We have indeed
f (2j+1) =
j∑
l=0
l∑
m=−l
F
(2j+1)
lm Y
(2j+1)
lm .
A Weyl map Ω2j+1 : C(S2) → M2j+1(C) is defined by simply mapping spherical harmonics
into spherical harmonics operators,
Ω2j+1 (Ylm (ϑ, ϕ)) =
{
Ŷ
(2j+1)
lm , l ≤ j,
0, l > j,
and extending the map by linearity. One can define the adjoint map as
Ω−1
2j+1
(
Ŷ
(2j+1)
lm
)
= Y
(2j+1)
lm (ϑ, ϕ).
Matrix Bases for Star Products: a Review 29
Using the Berezin symbol (8.3) it is possible to identify the adjoint map:
Ω†2j+1
(
F̂ (2j+1)
)
(ϑ, ϕ) = 〈ϑ, ϕ| F̂ (2j+1) |ϑ, ϕ〉 ,
for a matrix F̂ 2j+1. The two maps are one the adjoint of the other in the sense that
〈Ω2j+1(f), Ĝ(2j+1)〉2j+1 = 〈f,Ω†2j+1
(
Ĝ(2j+1)
)
〉L2(S2)
for all f ∈ C(S2) and all Ĝ(2j+1) ∈ M2j+1(C). The first scalar product is taken in the finite-
dimensional Hilbert space C2j+1, while the second one is taken in L2(S2).
The finite-dimensional matrix algebra is mapped by Ω2j+1 into a subspace of the infinite-
dimensional space of functions on S2. Restricting the functions of the sphere on this subspace
makes Ω†2j+1 = Ω−1
2j+1. This subspace is not an algebra under the usual commutative product of
functions, but it is a noncommutative algebra under the ?-product defined as usual by
(f ∗ g)(ϑ, ϕ) = Ω−1
2j+1(Ω2j+1(f)Ω2j+1(g)).
This ∗ product is given by the symbol of the product of two fuzzy harmonics [19, 41, 43], which
can be obtained in term of 6j-symbols [80]
Ŷ
(2j+1)
l′m′ Ŷ
(2j+1)
l′′m′′ =
j∑
l=0
(−1)2j+l
√
(2l′ + 1) (2l′′ + 1) (2j − l)(2j + l′ + 1)(2j + l′′ + 1)
4π(2j + l + 1)(2j + l′ + 1)(2j + l′′ + 1)
×
{
l′ l′′ l
j j j
}
C lml′m′l′′m′′ Ŷ
(2j+1)
lm . (8.5)
The fuzzy harmonics defined in (8.4) are the eigenvectors of the fuzzy Laplacian. The natural
infinitesimal action of SU(2) on M2j+1(C) is given by the adjoint action F̂ (2j+1) 7→ [Ji, F̂
(2j+1)]
of the generators Ji in the (2j+ 1)-dimensional representation. With these three derivations we
define the fuzzy Laplacian by the symbol of the operator
∇2 : M2j+1(C) 7→M2j+1(C),
∇2f (2j+1) = 2j+1〈ϑ, ϕ|∇2F̂ (2j+1)|ϑ, ϕ〉2j+1 = 2j+1〈ϑ, ϕ|
3∑
i=1
[
Ji,
[
Li, F̂
(2j+1)
]]
|ϑ, ϕ〉2j+1,
where, with an abuse of notation, we use the same symbol for the operator acting on M2j+1(C)
and the fuzzy Laplacian, which properly acts on the algebra of functions on the sphere C(S2).
Its spectrum consists of eigenvalues l(l + 1), where l = 0, . . . , 2j + 1, and every eigenvalue has
a multiplicity 2l + 1. The spectrum of the fuzzy Laplacian thus coincides up to order 2j + 1
with that of its continuum counterpart.
As in the case of the fuzzy torus described earlier the fuzzy sphere converges to the usual
sphere. Using properties of the 6j-symbols in the product (8.5) one can argue that the j →∞
limit of this product reproduces the standard product of spherical harmonics. This gives a naive
way to see that in the limit the fuzzy sphere algebra becomes the algebra of functions on S2.
In the sphere case there are rigorous proofs that this happens in a precise mathematical
sense [68]. The proof is based on the fact that the fuzzy sphere structure gives the algebra
of matrices a metric structure of a distance among states. It is possible also to prove that
the distance between the coherent states defined above converges to the metric distance on the
sphere [21]. Defining a distance among metric spaces makes it possible to show that the distance
between the fuzzy spheres and the ordinary sphere goes to zero as j →∞.
The fuzzy sphere as a matrix model has been studied extensively as a matrix model of field
theories, see for example the reviews [1, 63].
30 F. Lizzi and P. Vitale
8.3 The fuzzy disc
We have considered in Section 4.2 the matrix basis for the Wick–Voros product on the plane.
Let un now truncate the algebra R2
θ with the projector
P̂
(N)
θ =
N∑
n=0
|n〉〈n|.
The symbol of this operator is the function
P
(N)
θ (r, ϕ) =
N∑
n=0
〈z|n〉〈n|z〉 = e
r2
θ
N∑
n=0
r2n
θnn!
=
Γ
(
N + 1, r2/θ
)
Γ (N + 1)
, (8.6)
where we use the usual polar decomposition z = reiϕ. By construction P
(N)
θ ?V P
(N)
θ = P
(N)
θ .
The disc [48, 49, 50, 51] (see also [5]) is recovered considering the simultaneous limit
N →∞; θ → 0 with Nθ = R2, (8.7)
where R will be the radius of the disc. In the following we take R2 = 1 to simplify notations.
In this case the limit (8.6) can be performed using known properties of incomplete Gamma
functions to obtain
P
(N)
θ →
1 r < 1,
1/2 r = 1,
0 r > 1.
In other words, the symbols of the projector P̂
(N)
θ is an approximation of the characteristic
function of the disc, and converges to in the limit (8.7). This suggests to consider, in analogy
with the fuzzy sphere, a finite matrix algebra, ÂNθ (or rather a sequence of algebras), whose
symbols are functions with support on a disc. The fuzzy disc is thus defined as the sequence of
subalgebras ANθ ,
ANθ = P
(N)
θ ? R2
θ ? P
(N)
θ ,
with R2
θ the Wick Voros algebra on the plane.
A dual view, i.e. taking the projector I − P (N) gives a Moyal plane with a “defect” [66],
spherical wells have also been considered [71]. In order to do this we consider first the Laplacian
basis of functions for the disc with Dirichlet boundary conditions. For the Laplacian on the disc
all eigenvalues are negative, their modules λ are obtained solving for the zeroes of the Bessel
functions:
Jn
(√
λ
)
= 0.
They are doubly degenerate for n non zero, in which case they are simply degenerate. We label
them λn,k where k indicates that it is the kth zero of the function. The eigenfunctions are:
Φn,k = einϕ
(√
λ|n|,kr
2
)|n| ∞∑
s=0
(
−λ|n|,k
)s
s! (|n|+ s)!
(r
2
)2s
= einϕJ|n|
(√
λ|n|,kr
)
.
Because of relation (3.26) it is possible to express the Laplacian in terms of inner derivations,
and therefore, after the projection, express it as an automorphism of the algebra of matrices,
and find the eigenvectors of it. From the exact expression on the plane:
∇2f(z̄, z) = 4∂z̄∂zf =
4
θ2
[z, [f, z̄]?]?
Matrix Bases for Star Products: a Review 31
it is possible to define, in each A(N)
θ :
∇2
(N)f
(N)
θ = 〈z|∇2
(N)f̂
(N)
θ |z〉 ≡ 4
θ2
〈z|P̂ (N)
θ
[
â,
[
P̂
(N)
θ f̂ P̂
(N)
θ , â†
]]
P̂
(N)
θ |z〉.
The spectrum of this fuzzy Laplacian is of course finite, but it is possible to see from Fig. 1 that
it approaches the spectrum of the continuous Laplacian as N increases (with Nθ = 1).
y
140
120
100
80
60
40
x
252015105
20
y
140
120
100
80
60
40
x
252015105
20
y
140
120
100
80
60
40
x
252015105
20
Figure 1. Comparison of the first eigenvalues of the fuzzy Laplacian (circles) with those of the continuum
Laplacian (crosses) on the domain of functions with Dirichlet homogeneous boundary conditions. The
orders of truncation are N = 10, 20, 30.
The fuzzy Laplacian is an automorphism of the algebra on n × n matrices. In analogy
with the fuzzy harmonics described earlier we call the radial part of its eigenoperators Fuzzy
Bessel operators. Their symbols, the Fuzzy Bessel functions, form a basis for the fuzzy disc and
approximate well the actual Bessel functions, as can be seen from Fig. 2.
y
1
0,5
0
-0,5
-1
r
21,510,50
y
1
0,5
0
-0,5
-1
r
21,510,50
y
1
0,5
0
-0,5
-1
r
21,510,50
Figure 2. Comparison of the radial shape for the symbol Φ
(N)
0,1 (r, ϕ) (continuum line), the symbol of
the eigenmatrix of the fuzzy Laplacian for N = 10, 20, 30, with Φ0,1(r, ϕ).
Likewise it is possible [44], using the phase operator and phase states known in quantum
optics, to have fuzzy angles, i.e. some states concentrated in a small angular region of the disc.
Field theories on the fuzzy disc have been studied in [26, 47].
9 Matrix models and the emergence of gravity
Matrix models have a long and distinguished history, especially in string theory [7, 24, 42], in this
review we would like to discuss briefly how the discrete basis of the noncommutative products
describe earlier gives rise to a matrix model in which gravity is contemplated as an emergent
phenomenon, much like the emergent gravity of Sakharov [70]. Here by emergent gravity we
32 F. Lizzi and P. Vitale
really mean the emergence of fields moving in a curved background. This is a more modest
goal than having the metric degrees of freedom emerging as quantized fields. This would be
tantamount to have a full theory of quantum gravity. And as is known, we are not yet there . . . .
Since for the kind of products we are considering derivations are inner automorphisms of the
algebra2 since they can be expressed by a commutator:
∂
∂xµ
f = iθ−1
µν [xν , f ]∗. (9.1)
If one considers a U(1) gauge theory on this space, with unitary transformations given by star
unitary elements U ∗ U † = I, the action invariant for the transformation F → U ∗ F ∗ U † is
given by
S = −
∫
dx
1
4
F ∗ F.
One can define a covariant derivative
Dµf = ∂µf − i[f,Aµ]∗ = iθ−1
µν [Xµ, f ]∗
and
Fµν = [Dµ, Dν ]? = [Xµ, Xν ]? + θµν .
The connection between commutator with the coordinates and derivatives (9.1), suggest [55] the
definition of covariant coordinates
Xµ = xµ + θµνAν
and consequently
Dµf = iθ−1
µν [Xµ, f ]? = ∂µf − i[f,Aµ]?.
Therefore we have
Fµν = [Dµ, Dν ]? = [Xµ, Xν ]? + θµν .
The constant θ can be reabsorbed by a field redefinition and the action is the square of this
quantity, integrated over spacetime.
The action can therefore be rewritten, in the matrix basis as
S = − 1
4g
tr[Xµ, Xν ][Xµ′ , Xν′ ]gµµ′gνν′ , (9.2)
where the X’s are operators (matrices) and the metric gµµ′ is the flat Minkowski (or Euclidean)
metric.
We now briefly remind how gravity emerges from this model [75]. The equations of motion
corresponding to the action (9.2):
[Xµ, [Xν , Xµ′ ]]gµµ′ = 0.
2One would have to define precisely which algebra is being considered, since for example the coordinate
functions do not belong to the algebra of Schwarzian functions with the Moyal product. They however belong to
the multiplier algebra.
Matrix Bases for Star Products: a Review 33
These equations have different solutions, which we call vacua. One solution in particular cor-
responds to the star product generated by (3.10). We call the matrixes correspondding to this
particular solution X0, hence
[Xµ
0 , X
ν
0 ] = iθµν . (9.3)
Note that the relation can only be valid if the X’s are infinite matrices corresponding to non
bounded operators. Fluctuations around the X0’s will give a generalized commutation relation
[Xµ, Xν ] = iθ(X),
where we have defined the matrices Xµ = Xµ
0 +Aµ in analogy with the covariant coordinates.
An important result obtained in [75] (see also [83]) is obtained if one couples the theory to
a scaler field Σ. At this stage the meaning and origin of this field is yet undetermined, it is
a field which couples to the noncommutative space time. The free action of this field, using the
fact that the derivative are expressed as commutators with the coordinates, is:
tr[Xµ,Σ][Xν ,Σ]gµν ∼
∫
dx(Dµ′Σ)(Dν′Σ)θµµ
′
θνν
′
gµν =
∫
dx(DµΣ)(DνΣ)Gµν .
But it easy to recognize the fact that this is the action of field moving in a non flat background
described by the metric
Gµν(x) = θµµ
′
θνν
′
gµ′ν′ .
The mere fact that the field was moving in a noncommutative space described by the matrix
model has induce a curved background, so that gravity appears as an emergent phenomenon.
The vacuum (9.3) is not the only one. One can consider alternative vacua which have an
invariance for some group. For example
X̄µ
0 = Xµ
0 ⊗ In.
In this case the theory has an internal space, and a noncommutative U(n) symmetry. However
the U(1) degree of freedom of the theory is the described above, which couple gravitationally.
One can separate the trace part A0 form the the traceless generators of SU(n), considering as
fluctuations
X̄ = X̄0 +A0 +Aαλα.
Several models can be constructed based on these matrix models. Extra dimensions can appear
in the form of fuzzy spheres [17] and it is possible to have models which start having also
characteristics of the standard model [18, 35, 76].
Acknowledgements
We were partially supported by UniNA and Compagnia di San Paolo under the grant “Pro-
gramma STAR 2013”. F. Lizzi acknowledges support by CUR Generalitat de Catalunya under
project FPA2010-20807.
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http://arxiv.org/abs/1002.4937
http://dx.doi.org/10.1088/1126-6708/2000/05/010
http://arxiv.org/abs/hep-th/0003187
34 F. Lizzi and P. Vitale
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1 Introduction
2 The star product
3 The Weyl–Wigner and the s-ordered maps
3.1 The Moyal product
3.2 s-ordered symbols
3.3 The Wick–Voros product
3.4 Translation invariance
4 Matrix bases for s-ordered products
4.1 The Moyal matrix basis
4.2 The Wick–Voros matrix basis
5 Star products as reductions
6 Matrix basis for R3
7 Field theories on noncommutative spaces as matrix models
7.1 The Grosse–Wulkenhaar model
7.2 The translation invariant model
7.3 Gauge model on the Moyal plane
7.4 The scalar model on R3
7.4.1 The kinetic action in the fuzzy spherical harmonics base
8 Fuzzy spaces
8.1 The fuzzy torus
8.2 The fuzzy sphere
8.3 The fuzzy disc
9 Matrix models and the emergence of gravity
References
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