Introduction to renormalization
In these lectures I discuss peculiarities of the critical behaviour of “non-ideal” systems as it is explained by the renormalization group approach. Examples considered here include account of the single-ion anisotropy, structural disorder, frustrations. I introduce main ideas of renormalization a...
Saved in:
| Date: | 2006 |
|---|---|
| Main Author: | |
| Format: | Article |
| Language: | English |
| Published: |
Інститут фізики конденсованих систем НАН України
2006
|
| Series: | Condensed Matter Physics |
| Online Access: | http://dspace.nbuv.gov.ua/handle/123456789/121320 |
| Tags: |
Add Tag
No Tags, Be the first to tag this record!
|
| Journal Title: | Digital Library of Periodicals of National Academy of Sciences of Ukraine |
| Cite this: | Introduction to renormalization / Yu. Holovatch // Condensed Matter Physics. — 2006. — Т. 9, № 2(46). — С. 237–262. — Бібліогр.: 34 назв. — англ. |
Institution
Digital Library of Periodicals of National Academy of Sciences of Ukraine| id |
irk-123456789-121320 |
|---|---|
| record_format |
dspace |
| spelling |
irk-123456789-1213202017-06-15T03:03:05Z Introduction to renormalization Holovatch, Yu. In these lectures I discuss peculiarities of the critical behaviour of “non-ideal” systems as it is explained by the renormalization group approach. Examples considered here include account of the single-ion anisotropy, structural disorder, frustrations. I introduce main ideas of renormalization and show how it serves the explanation of typical features of criticality in the above systems: softening of the phase transition, changes in the universality class, complicated effective critical behaviour 2006 Article Introduction to renormalization / Yu. Holovatch // Condensed Matter Physics. — 2006. — Т. 9, № 2(46). — С. 237–262. — Бібліогр.: 34 назв. — англ. 1607-324X PACS: 05.50.+q, 05.70.Jk, 64.60.Ak DOI:10.5488/CMP.9.2.237 http://dspace.nbuv.gov.ua/handle/123456789/121320 en Condensed Matter Physics Інститут фізики конденсованих систем НАН України |
| institution |
Digital Library of Periodicals of National Academy of Sciences of Ukraine |
| collection |
DSpace DC |
| language |
English |
| description |
In these lectures I discuss peculiarities of the critical behaviour of “non-ideal” systems as it is explained by
the renormalization group approach. Examples considered here include account of the single-ion anisotropy,
structural disorder, frustrations. I introduce main ideas of renormalization and show how it serves the explanation
of typical features of criticality in the above systems: softening of the phase transition, changes in the
universality class, complicated effective critical behaviour |
| format |
Article |
| author |
Holovatch, Yu. |
| spellingShingle |
Holovatch, Yu. Introduction to renormalization Condensed Matter Physics |
| author_facet |
Holovatch, Yu. |
| author_sort |
Holovatch, Yu. |
| title |
Introduction to renormalization |
| title_short |
Introduction to renormalization |
| title_full |
Introduction to renormalization |
| title_fullStr |
Introduction to renormalization |
| title_full_unstemmed |
Introduction to renormalization |
| title_sort |
introduction to renormalization |
| publisher |
Інститут фізики конденсованих систем НАН України |
| publishDate |
2006 |
| url |
http://dspace.nbuv.gov.ua/handle/123456789/121320 |
| citation_txt |
Introduction to renormalization / Yu. Holovatch // Condensed Matter Physics. — 2006. — Т. 9, № 2(46). — С. 237–262. — Бібліогр.: 34 назв. — англ. |
| series |
Condensed Matter Physics |
| work_keys_str_mv |
AT holovatchyu introductiontorenormalization |
| first_indexed |
2025-07-08T19:38:57Z |
| last_indexed |
2025-07-08T19:38:57Z |
| _version_ |
1837108866070872064 |
| fulltext |
Condensed Matter Physics 2006, Vol. 9, No 2(46), pp. 237–262
Introduction to renormalization
Yu.Holovatch1,2,3
1 Institute for Condensed Matter Physics of the National Academy of Sciences of Ukraine,
1 Svientsitskii Str., 79011 Lviv, Ukraine
2 Institute für Theoretische Physik, Johannes Kepler Universität Linz, A–4040, Linz, Austria
3 Ivan Franko National University of Lviv, 79005 Lviv, Ukraine
Received February 13, 2006, in final form April 17, 2006
In these lectures I discuss peculiarities of the critical behaviour of “non-ideal” systems as it is explained by
the renormalization group approach. Examples considered here include account of the single-ion anisotropy,
structural disorder, frustrations. I introduce main ideas of renormalization and show how it serves the expla-
nation of typical features of criticality in the above systems: softening of the phase transition, changes in the
universality class, complicated effective critical behaviour.
Key words: critical behaviour, renormalization, field-theoretical renormalization group
PACS: 05.50.+q, 05.70.Jk, 64.60.Ak
Contents
1 Introduction 238
2 Criticality and scaling 238
3 Non-ideal spin Hamiltonians: single ion anisotropy, structural disorder, frustrations 240
3.1 An ideal system: m-vector model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 240
3.2 Single-ion anisotropy . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 240
3.3 Structural disorder . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 241
3.3.1 Random-site dilution . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 241
3.3.2 Random anisotropy . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 242
3.4 Frustrations . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 242
4 Renormalization 243
4.1 RG transformation . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 243
4.2 Fixed points and RG flows . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 245
5 From the spin Hamiltonians to the effective ones 247
5.1 m-vector model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 247
5.2 Cubic model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 249
5.3 Weakly diluted quenched m-vector model . . . . . . . . . . . . . . . . . . . . . . . . . . . . 249
5.4 Random anisotropy model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 251
5.5 Stacked triangular antiferromagnet . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 252
6 RG explanation of criticality in non-ideal systems 252
6.1 m-vector magnet . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 253
6.2 Cubic model . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 254
6.3 Random-site dilution . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 255
6.4 Random anisotropy . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 256
6.5 Stacked triangular antiferromagnet . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . . 257
7 Conclusions and outlook 258
c© Yu.Holovatch 237
Yu.Holovatch
1. Introduction
There exist many ways one can choose to approach the subject of these lectures. Taken that
the presentation is limited in time (spent by the students and lecturers at the Caribbean seashore
during the Mochima school in theoretical physics) and in space (given by the Editors of this
volume) this choice becomes a difficult one. The way of presentation I decided to follow was chosen
for several reasons. Extremely high theoretical level of certain modern renormalization group (RG)
studies of criticality in different systems (i.e. the “language” of these studies) sometimes does not
allow uninitiated reader to follow the derivations and even to understand the problem statement
and/or physical consequences of the results. Therefore the goal would be to provide a minimal
vocabulary, explaining main notions as simple as possible. On the other hand, it is tempting to
use such a simple vocabulary to make a short review of state-of-the art RG studies in a certain
domain. RG explanation of criticality in “non-ideal” 3d systems might be a good candidate for such
a domain: it is a subject of ongoing activity where important results have recently been obtained
and still a lot is to be done.
The following account will serve this purpose: after mentioning several examples of criticality
and scaling in condensed matter physics (section 2) and introducing model Hamiltonians of the
“non-ideal” systems we shall be interested in (section 3) I shall give the main ideas and notions
of the renormalization taking as an example a simple 1d Ising model (section 4). Once the reader
is acquainted with the RG transformation, its flow and fixed points, stability, universality and
scaling, I shall pass to the “non-ideal” systems showing how to obtain their effective Hamiltonians
and to the reviewing of recent results in this domain (sections 5 and 6). Some conclusions and
outlook are given in section 7.
2. Criticality and scaling
It is generally recognized that the term “critical point” was introduced in 1869 by Thomas An-
drews who studied a special point for carbon dioxide at about 31◦C and 73 atmospheres pressure
where the properties of liquid and of gas become indistinguishable [1]. Approaching the critical
point Tc from below, the liquid-gas density difference obeys a power law scaling governed by the
critical exponent β:
ρL − ρG ∝ (Tc − T )β, T → T−
c , (2.1)
as shown in the figure 1a. Similar power laws were found for some other physical quantities,
ρ
Tc
T
liquid
gas
(a)
M
Tc
T
(b)
Figure 1. Order parameters for the critical point of fluids (a) and Curie point of magnets (b).
describing their approach to zero or singularity at Tc. Less known is the fact, that the mean
field values of critical exponents (which are integers or their simple relations, e.g. β = 1/2 for
238
Introduction to renormalization
equation (2.1)) were questioned already over a century ago [2]1. Therefore much time passed
before there appeared the theory that both explained scaling behaviour (2.1) found in condensed
matter physics, and, on a larger scale, not only in physics, as well as offered quantitatively accurate
predictions for the exponents. As we now know this theory is based on the renormalization group
formalism [3] and has its origin in quantum field problems [4].
Next experimental observations of critical behaviour arrived in 1895, when Pierre Curie showed
that the ferromagnet iron also displayed a special point. This point is the highest temperature at
which iron can be permanently magnetized at zero external magnetic field which is now called the
Curie point. Curie himself noticed the parallelism between density-temperature curve at constant
pressure for carbon dioxide and magnetization-temperature curves at constant magnetic field for
iron. Behaviour of the order parameter, spontaneous magnetization M , in the vicinity of Curie
point (we denote it Tc as well) is governed by the familiar scaling law, cf. equation (2.1):
M ∝ (Tc − T )β, T → T−
c . (2.2)
It appears that different magnetic systems possess different sets of critical exponents but there
exist wide classes of magnets characterized by the same values of exponents. For example, uniaxial
magnets (like FeF2, MnF2, K2CoF4, Rb2CoF4) possess the same critical exponents. Moreover, they
coincide with those of simple fluids (Xe, SF6, CO2 . . . ) and some other systems. In such a case
it is said that these systems belong to the same universality class. The reasons for this as well as
the main characteristics of a system which define the universality classes will be cleared up in the
subsequent sections.
As far as the second derivatives of Gibbs free energy (isothermal magnetic susceptibility for
magnets and isothermal compressibility for fluids) diverge at Tc such phase transitions are referred
to as the second order phase transitions. This classification is due to Paul Ehrenfest. The order
parameter does not possess a discontinuity at Tc: it is an example of a continuous phase transition.
Phenomena which occur in the vicinity of the 2nd order phase transition point are called critical
phenomena. Changes of system structure at Tc are reflected in the behaviour of the pair correla-
tion function G2(r) (density-density function for fluids and magnetization-magnetization one for
magnets) and described by two universal exponents η and ν. The exponential decay of G2(r) with
distance at T > Tc transforms to the power law one:
G2(r) ∼ e−r/ξ , T > Tc ; G2(r) ∼
1
rd−2+η
, T = Tc , (2.3)
where ξ is a correlation length which diverges at Tc: ξ ∼ |T − Tc|−ν , T → Tc.
The XX century witnessed numerous experiments performed after pioneering works of Thomas
Andrews and Pierre Curie where it was shown that criticality and scaling accompany not only the
second order thermodynamic phase transitions. They are found in quantum phase transitions, per-
colation, non-equilibrium dissipative phase transitions. Properties of long flexible polymer chains
in good solvents are described in terms of critical phenomena as well. This list can be continued.
However already the above mentioned field of phenomena is very wide, especially if one takes into
account that any of them occurs in objects differing by their microscopic nature. Thermodynamic
2nd order phase transitions occur in magnets, ferroelectrics, quantum liquids; percolation phenom-
ena occurs in resistor networks, random magnets, gas filters. Nevertheless there are at least two
essential features common for the above mentioned phenomena. These are: the singular character
of change of properties in certain critical points and universal behaviour in the vicinity of these
points. As it became clear now the reason for both is anomalous growth of fluctuations and their
correlation at very high distances in the vicinity of critical points. Correlation range becomes the
only characteristic scale of the system and this causes the insensitivity of its behaviour to the
so-called microscopic parameters. In particular this results in the scale invariance of the system
where critical phenomena occur.
Since the pioneering work of Ising [5], theoretical description of many-particle systems is often
based on the so-called classical spin models. They appeared to be of primary importance to reveal
1I thank Reinhard Folk for pointing me the work of Jules-Émile Verschaffelt.
239
Yu.Holovatch
the main features of criticality. In the next section 3 we shall discuss several models currently used
in describing the critical phenomena, the analysis of which will be a subject of subsequent sections.
3. Non-ideal spin Hamiltonians: single ion anisotropy, str uctural disorder,
frustrations
In classical spin models, each particle is imitated as a vector located on a lattice site and
“interacting” with other vectors. The interaction is chosen to mimic the interparticle one whereas
the dimensionality of the vector is equal to the number of internal degrees of freedom of the particle.
Below we start with what we call an ideal system, which is a m-vector model and show how to
include different types of non-idealities within a spin model formalism.
3.1. An ideal system: m-vector model
The model describes a system of m-dimensional classical vectors (“spins”) located in the sites
of d-dimensional hypercubic lattice. The Hamiltonian of the m-vector model in the absence of an
external magnetic field reads:
H = −1
2
∑
R,R′
J (|R − R
′|) ~SR
~SR′ , (3.1)
where J(|R−R
′|) is an interaction between spins ~SR and ~SR′ located in sites R and R
′ (we shall
consider it to be the short-range one and of ferromagnetic origin: J(R) > 0) and ~SR
~SR′ means a
scalar product. Such a model appears as a natural generalization of the Ising model for the case
of m-component spin [6] and serves as a basic model in describing phase transitions in systems
with multicomponent order parameter. Below, we shall be mainly interested in the 3d systems.
However, for the sake of completeness let us recall that ferromagnetic ordering does not occur in
this model at d = 1 [6], it does not occur for d = 2 and m > 1 either [7]. Whereas the celebrated
Onsager solution [8] of the 2d Ising model (d = 2, m = 1) brings about the second order phase
transition with non-trivial values of the critical exponents: β = 1/8, ν = 1, η = 1/4, γ = 7/4,
α = 0 (exponents γ and α govern power law scaling of the magnetic susceptibility and specific
heat, the latter being logarithmically divergent for the 2d Ising model). At d = 3, ferromagnetic
ordering occurs at any m (see table 1 for typical values of exponents), and d = 4 is the upper
critical dimension of the problem: exponents attain their mean field values for d > 4.
Table 1. The standard values of the critical exponents of the 3d m-vector model (obtained in [9]
from the high-order renormalization group expansions).
m γ ν η β α
1 1.2396(13) 0.6304(13) 0.0335(250) 0.3258(14) 0.109(4)
2 1.3169(20) 0.6703(15) 0.0354(250) 0.3470(16) −0.011(4)
3 1.3895(50) 0.7073(35) 0.0355(250) 0.3662(25) −0.122(10)
We refer to the above model as to the ideal one: it describes the lattice system without any
defects of structure or complexities of interaction which are often encountered in real systems. Let
us show how those can be considered within the same formalism.
3.2. Single-ion anisotropy
Real substances often are anisotropic. For instance, in cubic crystals one expects the spin
interaction to react on the lattice structure (crystalline anisotropy) suggesting additional terms
in the Hamiltonian, invariant under the cubic group [10]. Such a single-ion anisotropy breaks the
240
Introduction to renormalization
rotational symmetry of the m-vector magnet (3.1) and the Hamiltonian reads:
H = −1
2
∑
R,R′
J (|R − R
′|) ~SR
~SR′ + V
∑
R
m
∑
i=1
(
Si
R
)4
, (3.2)
where V defines anisotropy strength and makes the order parameter to point either along the edges
(V > 0) or along diagonals (V < 0) of a m-dimensional hypercube. Therefore a model with the
spin Hamiltonian (3.2) is often called a cubic model.
An interesting phenomenon is observed in 3d cubic magnets: for low values of m their critical
exponents coincide with those of the “ideal” m-vector model (one speaks about isotropization of
critical fluctuations), whereas for large m they belong to the new universality class. The value
mcub
c at which one regime is changed to the other one is called the marginal dimension. Moreover,
transition to the low-temperature phase may also occur via the first-order scenario.
3.3. Structural disorder
To treat structural disorder within the lattice model one usually introduces random variables
into the Hamiltonian (3.1) and couples them to spin degrees of freedom. The new model can
either mimic a quenched system (new variables are randomly distributed and fixed in a certain
configuration) or an annealed, equilibrium one [11]. Here, we shall be interested in changes of critical
behaviour caused by quenched disorder. Furthermore, we shall consider two different examples,
showing how to introduce disorder via dilution and random anisotropy2.
3.3.1. Random-site dilution
To describe dilution, one may introduce the random-site m-vector model, considering situation,
when spins ~SR in (3.1) occupy only a part of the lattice sites, N1, the rest N−N1 sites being empty
(or occupied by non-magnetic atoms). Magnetic and non-magnetic sites are randomly distributed
and fixed in a certain configuration. The model Hamiltonian reads:
H = −1
2
∑
R,R′
J (|R − R
′|) cRcR′
~SR
~SR′ , (3.3)
where cR are the occupation numbers:
cR =
{
1, site R is occupied,
0, site R is empty.
(3.4)
Here we do not touch upon the phenomena occuring near the percolation threshold cperc and
consider the so-called weak dilution, c = N1/N � cperc. To complete the model one should choose
a certain distribution function for the occupation numbers cR. Let us consider the case when the
site is occupied with the probability c and is empty with the probability (1−c) and this probability
does not depend on the occupation numbers on the neighbouring sites. Such a situation corresponds
to the following occupation probability P({cR}):
P({cR}) =
∏
R
p(cR), p(cR) = cδ(cR − 1) + (1 − c)δ(cR). (3.5)
For different m, the above model describes magnetic phase transitions in crystalline alloys of uniax-
ial magnets and their non-magnetic isomorphs FexZn1−xF2, MnxZn1−xF2 [13], diluted Heisenberg-
like magnets [14,15], superfluid phase transition in He4 in porous medium [16]. Similar to the cubic
anisotropy systems, there exists a marginal dimensionality mdil
c separating two different scenarios:
2Another option would be to consider a random field disorder [12]. However, we do not introduce it since an
appropriate model differs from those considered here by its upper critical dimension and this will make our account
even broader. The same concerns strong dilution at the percolation threshold.
241
Yu.Holovatch
for m < mdil
c dilution causes a change in the universality class, whereas the asymptotic critical
exponents of m > mdil
c -component systems remain unchanged under dilution. The so-called Har-
ris criterion allows us to determine this marginal dimension from the pure system heat capacity
behaviour [17]. It states that critical exponents are not altered by dilution, if the heat capacity of
the pure system does not diverge (α < 0). From Table 1 one concludes that in three dimensions it
is the Ising model which changes the universality class.
3.3.2. Random anisotropy
Another way of introducing randomness to the model (3.1) is to consider the case, when each
spin is subjected to a local anisotropy of random orientation. The resulting Hamiltonian reads [18]:
H = −1
2
∑
R,R′
J(|R − R
′|)~SR
~SR′ − D
∑
R
(x̂R
~SR)2. (3.6)
Here, D > 0 is the strength of the anisotropy and x̂R is a random unit vector pointing in the direc-
tion of the local anisotropy axis. The random anisotropy model (3.6) is relevant to the description of
a wide class of disordered magnets. It was first introduced to describe magnetic properties of amor-
phous alloys of rare-earth compounds with aspherical electron distributions and transition metals
[18]. Today the majority of the amorphous alloys containing rare-earth elements are recognized to
be random anisotropy magnets [19].
As in the random-site case, the model should be completed by choosing a certain distribution
p(x̂R) for the random variables x̂R. Most often, two different distributions are considered [21]. The
first is an isotropic one, where the random vector x̂R points with equal probability in any direction
in the m-dimensional hyperspace:
p(x̂R) ≡
(∫
dmx̂R
)−1
=
Γ(m/2)
2πm/2
. (3.7)
Here Γ(x) is Euler gamma-function, and the right-hand side presents the volume of the m-
dimensional hypersphere of unit radius. This distribution mimics an amorphous system without any
preferred direction. The second distribution restricts the vector x̂R to point with equal probability
along one of the 2m directions of the axes k̂i of a hypercubic lattice:
p(x̂R) =
1
2m
m
∑
i=1
[
δ(m)(x̂R − k̂i) + δ(m)(x̂R + k̂i)
]
, (3.8)
where δ(y) are Dirac δ-functions. This distribution (sometimes called a cubic one) corresponds to
a situation when an amorphous magnet still remembers the initial cubic lattice structure.
It is generally believed that ferromagnetism does not exist for the 3d model (3.6) with the
isotropic random axis distribution (3.7) [19]. However, anisotropic distribution (3.8) leads to a
magnetically ordered low temperature phase and the transition belongs to the random site Ising
model universality class [20].
3.4. Frustrations
An archetype of a model describing the effect of frustrations is the model of stacked triangular
antiferromagnet [22]. In this model, the antiferromagnetically interacting spins are placed on the
sites of 2d triangular lattices stacked in register along the orthogonal direction. The sign of the
interlayer interaction is unimportant, because there are no frustrations in orthogonal direction.
The Hamiltonian reads:
H = −1
2
∑
〈R,R′〉
J ~SR
~SR′ , J =
{
J1 < 0, inside a plane,
J2 , between planes.
(3.9)
242
Introduction to renormalization
Sum in (3.9) runs over the nearest neighbours of the above described lattice. Systems which are
characterized by the Hamiltonian (3.9) exhibit noncollinear spin ordering. An example is given by
the famous 120◦ structure: each spin in a layer forms 120◦ angles with the neighbouring spins.
Although model (3.9) is formulated for general m, of most interest are values m = 2 and m =
3. Namely for these values of m the model has experimental realizations [14,23] and describes
noncollinear ordering of stacked triangular antiferromagnets as VCl2, VBr2, CsMnBr3, and helical
magnets as Ho, Dy, β−MnO2. In the latter substances, frustration is induced by the competition of
ferromagnetic nearest-neighbour and antiferromagnetic next-nearest-neighbour interactions, which
acts only along one lattice axis: J1 > 0 and J2 < 0, correspondingly. When the ratio J1/J2 exceeds
a critical value, in the low-temperature phase the spins align ferromagnetically in a plane and
form a spiral along the orthogonal axis. Moreover, at m = 3 Hamiltonian (3.9) also describes A/B
transition in He3.
Numerous experimental and MC studies performed so far have not lead to the definitive con-
clusion about the order of transition into non-collinear state. There is no unique answer from the
theoretical viewpoint either [14,23].
4. Renormalization
In order to explain the main ideas of the renormalization group (RG) theory and to show how
this method works in practice we shall study the critical behaviour of the 1d Ising model by means
of the RG approach. The explanations given below are due to Michael Fisher [24], who compared
this study with the Bohr-Sommerfeld picture in quantum mechanics. Passing from the classical
mechanics to the full account of quantum mechanics, Bohr-Sommerfeld’s picture represents only a
crude approximation. Nevertheless, it introduces some important ideas, just like RG study of the
1d Ising model enables one to introduce ideas of renormalization and scaling in the critical region.
4.1. RG transformation
We start from the Hamiltonian of the 1d Ising model in the presence of an external magnetic
field H , which being divided by kBT is written in the following form:
−Heff = K
N
∑
j=1
SjSj+1 + h
N
∑
j=1
Sj + C
N
∑
j=1
1, (4.1)
where K = J/(kBT ), h = H/(kBT ), and the last term C
∑N
j=1 1 = NC is added for the convenience
of forthcoming calculations. We call this temperature-dependent Hamiltonian an effective one, Heff .
Note, however, that contrary to the effective Hamiltonians considered in the forthcoming section 5,
it contains full information about the spin Hamiltonian. Given {K, h, C} one completely specifies
Heff : thus it can be regarded as a point in a space of 3 parameters {K, h, C}. With change of
{K, h, C}, this point moves.
One of the first approaches to the RG is to regard it as a specific way of calculating the partition
function ZN{H}:
ZN{Heff} =
1
2N
∑
S1=±1
· · ·
∑
SN=±1
e−Heff , (4.2)
where we have normalized the expression for the partition function by the partition function of
the ideal model (Z ideal = 2N ) to obtain ZN{Heff} →T→∞ 1. The RG method of calculating ZN
is to step-by-step perform the summation in (4.2) in such a way as to try as much as possible to
preserve the system as it used to be prior the summation.
Let us do this by means of decimation procedure: performing summation over every second
spin variable (see figure 2). To show the changes which are introduced by summation over certain
spin variable S0 let us represent the total Boltzmann weight in the factorized form:
e−Heff = · · · eKS−S0+
1
2h(S−+S0)+CeKS0S++ 1
2h(S0+S+)+C · · · = · · ·P (S−, S0)P (S0, S+) · · · (4.3)
243
Yu.Holovatch
Figure 2. Decimation procedure followed by the spatial rescaling. Summation over every second
spin variable is performed and then all lengths are rescaled in such a way that the new lattice
spacing a′ is equal to the old one a. Now any distance R in the original lattice becomes R′ = 1/2R
in the new one measured in the units of the lattice spacing.
Summation over S0 which enters only P (S−, S0) and P (S0, S+) will lead to a new function
P ′(S−, S+), defined by:
P ′(S−, S+) =
1
2
∑
S0=±1
P (S−, S0)P (S0, S+). (4.4)
Factor 1/2 is included in (4.4) because with each spin eliminated one must remove factor 1
2 from
the normalizing factor in (4.2). Now the RG idea is to express the new factor P ′ in the same form
as the initial one:
P ′(S−, S+) = eK′S−S++ 1
2h′(S−+S+)+C′
. (4.5)
The new parameters define the renormalized Hamiltonian:
H′
eff = H′
eff{K ′, h′, C′}, (4.6)
and H′
eff has a half of initial spins. The result can be formally written as :
H′
eff = Rb[Heff ], (4.7)
with a spatial rescaling factor b (in our case b = 2, see figure 2). The resulting number of spins
N ′ is connected with the initial one by N ′ = N/b, and in the case of d-dimensional system this is
generalized to:
N ′ = N/bd. (4.8)
Performing summation (4.4) and expressing the result in the renormalized form (4.5) one can
get expressions for the renormalized variables K ′, h′, C′ (it is proposed to the interested reader to
do this).
Now the renormalized model is characterized by the Hamiltonian H′
eff(K ′, h′, C′) having the
form similar to the initial one, though the lattice constant after taking the sum over every second
spin is equal to 2a (see figure 2). In order to have the renormalized model looking as the initial one
we rescale all lengths in such a way that the new lattice spacing a′ is equal to the old one. Now
any distance R in the original lattice becomes R′ = 1/2R in the new one measured in the units of
the lattice spacing. For the arbitrary b we have the following mapping:
R ⇒ R′ = R/b. (4.9)
244
Introduction to renormalization
Now let us have a look at the behaviour of the spin-spin correlation function 〈S0SR〉. First let us
renumber the remaining spins to have the labels arranged in the subsequent order (see figure 2):
S2 ⇒ S′
1, S4 ⇒ S′
2, . . . , S2k ⇒ S′
k, . . . Or, taking into account (4.9), S2R′ = S′
R′ . Since
the remaining spins S′ have not changed under the renormalization procedure, the renormalized
correlation function is equal to the original one:
〈S0S2R′〉 = 〈S′
0S
′
R′〉. (4.10)
It follows that if the original correlation length is ξ = ξ(Heff) then the renormalized correlation
length is two times smaller. Or, for a general b:
ξ(Heff) = bξ(H′
eff). (4.11)
The RG procedure has the effect of shrinking the correlation length. Recalling that ξ becomes
infinite at T = Tc one can state, that the RG procedure is driving a system away from criticality
(if it was not critical). Already here it is seen that the RG transformation we are considering has
a deep connection with the critical properties of a system.
4.2. Fixed points and RG flows
To proceed in the determination of these properties let us have a closer look at the partition
function and at the free energy of the system. First, after taking a partial trace of the partition
function ZN{Heff} we get the Boltzmann distribution in N ′ spin variables S′:
e−H′
eff (S
′) = SpS′′
N ′′e−Heff (S), (4.12)
where SpS′′
N ′′ stands for the trace over N ′′ = N − N ′ spin variables S′′. Now taking trace over the
remaining N ′ spin variables one gets:
ZN ′{H′
eff} = SpS′
N ′e−H′
eff (S
′). (4.13)
Substituting into the right hand side of (4.13) its explicit form given by (4.12) one gets:
ZN ′{H′
eff} = SpS′
N ′SpS′′
N ′′e−Heff (S) = ZN{Heff}.
This relation can be rewritten in terms of the flow equations which describe a motion of a point
describing an effective Hamiltonian:
K ′ = RK(K, h), (4.14)
h′ = Rh(K, h), (4.15)
C′ = bdC + R0(K, h), (4.16)
for the couplings defining Heff . As far as the temperature T enters parameter K, the above relations
imply the flow equations for T as well. To study this let us consider H = 0 and write the flow
equation for the temperature as:
T ⇒ T ′ = R(T ), (4.17)
with R(T ) being appropriate function of T . Suppose that R(T ) has a form given in the figure 3.
The important feature there is that R(T ) crosses the line T ′ = T at some point T = T ∗. This point
is called the fixed point (FP). It is clear that when T is smaller or greater than T ∗ the successive
application of the renormalization procedure drives the system away from the fixed point. On the
other hand at T = T ∗ the system remains at the fixed point under the application of the RG
procedure. Recalling the flow equation for the correlation length (4.11) which can be written as:
ξ(T ) = bξ(T ′), (4.18)
one has that in the FP this reads:
ξ(T ∗) = bξ(T ∗). (4.19)
Since b > 1 it is possible only when
245
Yu.Holovatch
Figure 3. Temperature renormalization function R(T ). R(T ) crosses the line T ′ = T at the
fixed point T = T ∗. When T is smaller (T = T0) or greater (T = T1) than T ∗, the successive
application of the renormalization procedure drives the system away from the fixed point. On
the other hand at T = T ∗ the system remains at the fixed point under the application of the
RG procedure.
• ξ(T ∗) = ∞ and
• ξ(T ∗) = 0.
The first case characterizes the critical point T ∗ = Tc, whereas the second corresponds to the
vanishing of the correlation length at zero or infinite temperatures (when the spins are frozen at a
ground state or totally uncoupled). Note, that this corresponds to two FPs more (T ∗ = 0, T ∗ = ∞)
in the figure 3.
Having identified the fixed point T ∗ of the RG transformation as the critical point of the system
let us study what knowledge about the critical exponents can be obtained based on the properties
of the RG transformation. First let us linearize it in the vicinity of the FP introducing the variable:
τ =
T − Tc
Tc
=
T − T ∗
T ∗
(4.20)
and replacing the plot of R(T ) near T ∗ by its tangent at T ∗.
Then after the renormalization the temperature deviation will be:
τ ′ ≡ τ (1) ' Λ1(b)τ (4.21)
for small enough τ , where Λ1(b) is the slope of the tangent. To find its dependence on b let us
apply the renormalization procedure twice:
τ ′′ ≡ τ (2) ' Λ1(b)Λ1(b)τ. (4.22)
Such transformation should be equivalent to transforming with a spatial rescaling factor b2:
τ (2) ' Λ1(b
2)τ, (4.23)
or
Λ1(b)Λ1(b) = Λ1(b
2). (4.24)
This leads to the conclusion that Λ1(b) should have the following form:
Λ1(b) = bλ1 , (4.25)
246
Introduction to renormalization
with λ1 being constant independent of b. Now for the correlation length after l transformations
one gets:
ξ(τ) = blξ([Λ1(b)]
lτ) = blξ(blλ1τ). (4.26)
As far as (4.26) holds for any l (and for small τ(!)) let us choose it to satisfy:
blλ1τ = 1,
or
bl = (1/τ)1/λ1 . (4.27)
Then from (4.26) one gets:
ξ(τ) = (1/τ)1/λ1ξ(1)
or
ξ(τ) ∼ τ−1/λ1 . (4.28)
Comparing (4.28) with the definition of the critical exponent ν: ξ(τ) ∼ τ−ν we get:
ν = 1/λ1. (4.29)
So as we have seen, the knowledge of the linearized RG transformation enables one to determine
the critical exponent!
Note, that having carried out similar analysis for the free energy per spin, f = −N−1 lnZN ,
one arrives at the expression: f(τ) = b−dlf(bλ1lτ). Again, choosing b from (4.27) one gets f(τ) =
τd/λ1f(1) = τdνf(1). Comparing the latter with the scaling behaviour of the free energy in the
vicinity of the critical point f(τ) ∼ τ2−α one proves the hyperscaling relation:
2 − α = dν. (4.30)
Similarly, considering the case of non-zero magnetic field one can write the recursion relations
for the renormalized temperature and field. Their linearization in the vicinity of the fixed point
leads to the matrix of the linear RG operator with two eigenvalues, defining two different critical
exponents.
To complete the RG study of the 1d Ising model it remains to perform the above described
procedure explicitly and to find the values of the critical exponents (at zero-temperature FP T =
H = 0, [24]). The above described transformations concisely give the main features of the RG
transformation, allowing us, in particular, to define the critical exponents of the system.
5. From the spin Hamiltonians to the effective ones
Very often a starting point for the RG study of critical behaviour of a many-particle system
is its effective Hamiltonian. Taken, that in the model description the system is identified with the
spin Hamiltonian (see section 3), the effective Hamiltonian arises as a certain “metamodel”. It
shares global features of different spin Hamiltonians: their dimensionalities, symmetries, type of
interparticle interaction and as a result, brings about common features of their critical behaviour.
Below, we shall show how to obtain the effective Hamiltonians for the spin models of section 3.
5.1. m-vector model
To get an effective Hamiltonian we proceed as follows. Let us define the free energy F and the
partition function Z of a spin model (3.1) as:
F = −β−1 lnZ, Z = Spe−βH, (5.1)
where β = (kBT )−1 and Sp(. . . ) as usually means the sum over all possible states. In our case it
corresponds to the integration over the surface of m-dimensional hypersphere (we take it to be of
unit radius):
Sp(. . . ) =
∏
R
∫
d~SRδ(|~SR| − 1)(. . . ). (5.2)
247
Yu.Holovatch
Let us introduce the Fourier-transforms of the variables ~SR by:
~SR =
1√
N
∑
k
eikR~Sk , ~Sk =
1√
N
∑
R
e−ikR~SR . (5.3)
Here and below, when it will not be mentioned explicitly, vector R spans all sites of the lattice
whereas k changes in the first Brillouin zone. Now the Hamiltonian (3.1) can be rewritten as :
H = −1
2
∑
k
ν(k)~Sk
~S−k , (5.4)
where we have introduced Fourier transform of the potential of interaction by:
J(R) =
1
N
∑
k
eikRν(k), ν(k) =
∑
R
e−ikRJ(R). (5.5)
Now the partition function (5.1) reads:
Z = Spe
β
2
�
k
ν(k)~Sk
~S−k . (5.6)
In order to take the trace in (5.6) let us transform it to the expression with linear dependence on
~Sk. To this end, one makes use of the Stratonovich-Hubbard transformation introducing the field
variable ~φk which is conjugated to the spin variable ~Sk by the identity:
e−βH =
∏
~k
e
βν(k)
2
~Sk
~S−k =
∏
~k
(
1
2πβν(k)
)m/2
∫
d~φke
−1
2βν(k)
~φk
~φ−k+~Sk
~φ−k
∼
∫
(d~φ)e
�
k
( −1
2βν(k)
|~φk|
2+~Sk
~φ−k). (5.7)
As far as the Stratonovich-Hubbard transformation was performed for each function ~Sk we arrived
at the functional integral
∫
(d~φ) =
∏
R
∏m
i=1
∫∞
−∞ dφ
(i)
R
over the field variables
~φR =
1√
N
∑
k
eikR~φk , ~φk =
1√
N
∑
R
e−ikR~φR . (5.8)
From now on we omit the coefficients in front of the functional integral.
We have reached our goal: now the trace in the expression for the partition function concerns
only the last term of the exponents under integration in (5.6) which is a linear function of ~S:
Z = Spe
β
2
�
k
ν(k)~Sk
~S−k ∼
∫
(d~φ)e
−1
2
�
k
1
βν(k)
|~φk|
2
Sp e
�
R
~SR
~φR . (5.9)
The last step is to take integral over ~SR (recall that the trace is defined by equation (5.2)). This
can be achieved by passing from the m-dimensional Cartesian coordinates S1, S2, . . . , Sm to the
m-dimensional polar ones S, θ1, . . . , θm−1. Performing this integration and representing the result
as a series in ~φ one gets for the partition function:
Z ∼
∫
(d~φ)e
−1
2
�
k
( 1
βν(k)
−u2)|~φk|
2−
�
R
�
∞
l=2
u2l
(2l)!
|~φR|2l
, (5.10)
where the coefficients u2l readily follow: u2 = −1/m, u4 = 6/m2(m + 2), . . . Expression (5.10)
gives the functional representation of the partition function of the m-vector model. So far we have
not gained a lot by the above described transformations: the difficulty of taking Sp of the initial
expression (5.4) now is transformed into the difficulty of calculating a functional integral (5.10).
However, in order to study critical behaviour of the model (3.1), expression (5.10) can be further
approximated. It appears that all of φ higher than the fourth powers do not effect the asymptotic
248
Introduction to renormalization
critical behaviour at d = 3: they do not change the value of the fixed point of RG transformation
(cf. section 4) and are irrelevant in the RG sense [3,4]. Therefore one can be restricted to the φ4
model. Being interested in the long-range correlations arising in the system in the vicinity of a
critical point, one substitutes the Fourier image of the interaction potential by its expansion for
small wave vector values k: ν(k) ' ν(0)−1/2|ν′′(0)|k2 and writes the contribution
∑
k
k2~φk
~φ−k in
the form
∑
R
(∇~φR)2 re-scaling variables φ to get the gradient term without any coefficient. One
further passes to the continuous limit
∑
R
→
∫
dR and gets for the free energy:
F ∼ ln
∫
(d~φ)e−Heff (5.11)
with the effective Hamiltonian:
Heff =
∫
ddR
{
1
2
(
(∇φ)2 + µ2
0φ
2
)
+
u0
4!
φ4
}
, (5.12)
where φ ≡ ~φR, φ2 ≡ |~φR|2, variables µ0 and u0 are called bare mass and coupling. From the above
derivation we know that u0 is positive (u0 ∼ u4 from equation (5.10)), which ensures the existence
of the integral (5.11). Effective Hamiltonian (5.12) shares common global properties with the spin
Hamiltonian (3.1): the dimension of space (i.e. dimension of vectors k, R), dimension of the order
parameter (m) and its symmetry: the functional representation we obtained is symmetric under
group of rotations in the m-dimensional space O(m): it depends only on |~φ|. Let us now see what
differences will appear in the functional representation of the non-ideal models.
5.2. Cubic model
It is intuitively clear that an effective Hamiltonian of the cubic model (3.2) will differ from that
of the m-vector model (5.12) by the presence of terms with new, cubic symmetry. Indeed this is the
case that one can easily check following the next steps: starting from the Hamiltonian (3.2) via the
Stratonovich-Hubbard transformation one obtains the functional representation for the interaction
part (which coincides with equation (5.7)). An additional term e−βV
�
i
Si
can be represented in
a form of the functional derivative e−βV
�
i
∂4/∂(φi)4 , resulting in the following expression for the
partition function of the cubic model (c.f. equation (5.9)):
Z ∼
∫
(d~φ)e
−1
2
�
k
1
βν(k) |
~φk|
2
e−βV
�
R
�
i
∂4/∂(φi
R
)4Sp e
�
R
~SR
~φR . (5.13)
Taking trace in equation (5.13) leads to the familiar expression (5.10). However, the derivative
∂4/∂(φi
R
)4 gives rise to the contributions of cubic symmetry, the lowest order contribution being
proportional to v0
∑
i(φ
i
R
)4. The resulting effective Hamiltonian reads:
Heff =
∫
ddR
{
1
2
(
(∇φ)2 + µ2
0φ
2
)
+
u0
4!
φ4 +
v0
4!
m
∑
i=1
(φi)4
}
, (5.14)
and contains two bare couplings u0 and v0, corresponding to the φ4 terms of different symme-
tries. Coupling u0 is positive, whereas the sign of the coupling v0 coincides with that of V in
equation (3.2). At v0 = 0 one recovers an effective Hamiltonian of the m-vector model (5.12).
5.3. Weakly diluted quenched m-vector model
The peculiarities of the free energy calculation for the model (3.3) consist in averaging over
quenched disorder. Indeed, for each configuration of empty and occupied sites in (3.3) one can
write a corresponding configuration-dependent partition function Zconf :
Zconf = Sp~Se−βH, (5.15)
249
Yu.Holovatch
where Sp~S concerns spin degrees of freedom and is defined by (5.2). The free energy is obtained as
the configurational average:
F = −β−1〈lnZconf〉conf . (5.16)
One of the ways to proceed is to make use of the replica trick [25], which allows us to avoid
integration of the logarithm in (5.16) substituting it by a power function:
lnZ = lim
n→0
Zn − 1
n
. (5.17)
Then Zn
conf can be written as:
Zn
conf = Spe
β
2
�
R,R′ J(|R−R
′|)
�
n
α=1 ~σα
R
~σα
R′ . (5.18)
with obvious notations Sp(. . . ) =
∏n
α=1 Sp~Sα(. . . ) and ~σα
R
≡ cR~Sα
R
. Upcoming calculations closely
follow lines of the subsection 5.1. Introducing by the Stratonovich-Hubbard transformation field
variables ~φα
R
, conjugated to ~σα
R
one is able to take the trace over spin subsystem and is left with
the configuration-dependent partition function:
Zn
conf ∼
∫
(d~φ)e
−1
2
�
k
1
βν(k)
�
n
α=1
~φα
k
~φα
−k
−
�
R
�
∞
l=1
u2l
(2l)!
�
n
α=1 |~φα
R
|2lcR . (5.19)
Note that cR appears in (5.19) in the first power because from cR = {0, 1} it follows that (cR)l ≡
cR. The last step is to perform configurational averaging of (5.19) with the distribution function
(3.5). Let us represent the result of averaging in the exponential form
〈
e−
�
R
�
∞
l=1
u2l
(2l)!
�n
α=1 |~φα
R
|2lcR
〉
conf
=
∏
R
e
�
p>1
(−1)p
p! κp(c)
(�
∞
l=1
u2l
(2l)!
�
n
α=1 |~φα
R
|2l
)p
, (5.20)
where κp(c) are cumulants of random variables cR and can be easily calculated for the random
variable distribution (3.5):
κ1 =
∑
cR={0,1}
cRp(cR) = c,
κ2 =
∑
cR={0,1}
(cR)2p(cR) −
∑
cR={0,1}
cRp(cR)
2
= c(1 − c), . . . . (5.21)
For the free energy one gets:
F = −β−1〈lnZconf〉conf = −β−1 lim
n→0
{
c1
n
∫
(d~φ)ne−F [φ] − 1/n
}
, (5.22)
where the (n-dependent) coefficient c1 can be recast explicitly following all the steps of calculations
described above and the free energy functional F [φ] is given by:
F [φ] =
1
2
∑
k
n
∑
α=1
1
βν(k)
|~φα
k
|2 −
∑
p>1
(−1)p
p!
κp(c)
∑
R
(
∞
∑
l=1
u2l
(2l)!
n
∑
α=1
|~φα
R
|2l
)p
. (5.23)
Again, as in the previous subsection, restricting ourselves to the φ4 terms, expanding the short-
range interaction potential, re-scaling the fields and passing to the continuous limit one gets for
the free energy:
F ∼
∫
(d~φ)e−Heff , (5.24)
250
Introduction to renormalization
now the proportionality sign also hides, besides the coefficient, the replica limit, as written explicitly
in (5.22). The effective Hamiltonian reads:
Heff =
∫
ddR
1
2
n
∑
α=1
[
(∇φα)2 + µ2
0(φ
α)2
]
+
u0
4!
n
∑
α=1
(φα)4 +
v0
4!
n
∑
α,β=1
(φα)2(φβ)2
. (5.25)
Here, the coupling u0 is positive (being proportional to cu4, (5.10)) whereas the coupling v0 is
proportional to c(c − 1)u2
2 and is negative. The last term in (5.25) is present only for non-zero
dilution: it is directly responsible for the effective interaction between replicas due to the presence
of impurities.
5.4. Random anisotropy model
To treat the random anisotropy in the spin Hamiltonian (3.6), one first writes the configuration-
dependent partition function for a fixed local anisotropy axes configuration {x̂}:
Zconf({x̂}) = Sp~Se−βH. (5.26)
As in equation (5.15), trace in equation (5.26) concerns only the spin degrees of freedom. Applying
the Stratonovich-Hubbard transformation to the interaction part of the Hamiltonian one presents
(5.26) in the form of the functional integral and is able to take the trace:
Zconf ({x̂}) ∼
∫
(d~φ)e−H(x̂), (5.27)
with
H(x̂) =
∫
ddR
{1
2
[
(∇φ)2 + µ2
1φ
2
]
− D1(φx̂)2 + v1φ
4 + z1φ
2(φx̂)2 + · · ·
}
. (5.28)
Again, the expansion for small k was performed and the continuous limit has been taken. Explicit
values for the coefficients µ1, D1, v1, z1 are given in [20].
The rest of calculations follow the steps outlined in the section 5.3: for the quenched disorder
the free energy is defined by (5.16), where averaging is performed over the random axis distribution
(given by (3.7) or (3.8)). Subsequently, one substitutes lnZ by a power function via the replica
trick (5.17). For the isotropic random axis distribution (3.7) one gets the effective Hamiltonian [21]
containing three φ4 couplings of different symmetry:
Heff =
∫
ddR
1
2
[
(∇ϕ)2 + µ2
0ϕ
2
]
+ u0ϕ
4 + v0
n
∑
α=1
(φα)4 + w0
n
∑
α,β=1
m
∑
i,j=1
φα
i φα
j φβ
i φβ
j
. (5.29)
Here and below ϕ2 =
∑n
α=1(φ
α)2. One can check the signs of the couplings: u0 ∼ D2
m2(m+2) > 0,
v0 ∼ D > 0, w0 ∼ −D2
m(m+2) < 0. Moreover, from the explicit expressions for the couplings w0 and u0
one gets for their ratio w0/u0 = −m. The latter relation determines a region of physically allowed
initial values in the (u − v − w)-space of couplings.
For the cubic distribution (3.8) the average over the random variables {x̂} leads to the effective
Hamiltonian [21] with four couplings:
Heff =
∫
ddR
1
2
[
(∇ϕ)2 + µ2
0ϕ
2
]
+ u0ϕ
4 + v0
n
∑
α=1
(φα)4
+w0
m
∑
i=1
n
∑
α,β=1
(φα
i )
2
(
φβ
i
)2
+ y0
m
∑
i=1
n
∑
α=1
(φα
i )
4
. (5.30)
251
Yu.Holovatch
The signs of the couplings are as follows: u0 ∼ D2/(2m2) > 0, v0 ∼ D > 0, w0 ∼ −D2/(2m) < 0.
The last term in (5.30) is of cubic symmetry. It has to be included since it is generated if one further
applies the RG transformation. Therefore y0 can be of either sign. The symmetries of w0 terms in
(5.29) and (5.30) differ. However the ratio w0/u0 = −m holds for the effective Hamiltonian (5.30)
as well.
5.5. Stacked triangular antiferromagnet
It is an interaction potential (3.9) that makes a difference in a derivation of the effective
Hamiltonian for the frustrated magnet and the regular one. One can find the details of the procedure
to derive an effective Hamiltonian via the Stratonovich-Hubbard transformation in [22]. The main
difference from the above described calculations for the m-vector magnet (subsection 5.1) arises
when one proceeds with the Gaussian term in equation (5.10):
∑
k
(
1
βν(k)
− u2
)
|~φk|2 ≡
∑
k
d(k)|~φk|2. (5.31)
The sum over k in (5.31) spans the 1st Brillouin zone. For the m-vector model, d(k) has a minimum
at k = 0 and an expansion in k around this minimum results in the effective Hamiltonian (5.12).
Note, that one minimum in d(k) leads to one field variable φk in the effective Hamiltonian. For
the frustrated model, the d(k) has a maximum at k = 0 and two independent minima at k 6= 0 in
the 1st Brillouin zone. Subsequently, the zone can be rearranged into separate subzones and the
field φk over the original Brillouin zone can be decomposed into two fields, φ1,k and φ2,k, with k
confined to the new subzone. Now, the minima of d(k) occur at the origin and pertain to the fields
φ1 and φ2. The above decomposition results in the appearance of the terms of new symmetry in
the effective Hamiltonian, which now reads:
Heff =
∫
ddR
{
1
2
[
µ2
0(φ
2
1 + φ2
2) + (∇φ1)
2 + (∇φ2)
2
]
+
u0
4!
[
φ2
1 + φ2
2
]2
+
v0
4!
[
(φ1 · φ2)
2 − φ2
1φ
2
2
]
}
,
(5.32)
with a scalar product of m-component fields φ1 · φ2 =
∑m
i=1 φi
1φ
i
2. The coupling u0 in (5.32) is
positive, whereas the sign of the coupling v0 determines a type of spin ordering: a non-collinear
order occurs for v0 > 0. For v0 < 0 the fields φ1 and φ2 tend to allign parallel or antiparallel, this
corresponds to the sinusoidal ordering or the linearly-polarized spin-density wave.
6. RG explanation of criticality in non-ideal systems
Now with the effective Hamiltonians (section 5) and the RG method (section 4) at hand one
can proceed further, applying the method in order to study the critical behaviour of the models.
However, to be honest with the reader, he should be warned that the whole story is not that
simple as shown in the section 4. It is not only because the very models we are interested in are
much more complicated as compared to the 1d Ising model: in what follows below we will be
interested in the criticality of 3d systems. The real space renormalization described in the section 4
works for the low-dimensional systems, whereas for realistic 3d systems one rather applies the RG
equations in the momentum space. Moreover, the very approach to RG as to the way of calculating
the partition function of the system [3] has its alternative known as the field-theoretical RG [4].
The latter is a tool to cope with the divergencies of correlation functions (vertex functions). Here,
the RG procedure consists in the controlled rearrangement of the perturbation theory expansions
giving rise to the RG equations. However, the underlying notions of both procedures are the same:
given the effective Hamiltonian one applies the RG transformation and studies the flow equations
for the couplings of the Hamiltonian. If the flow equations possess the fixed point (FP), which is
stable and reachable from the initial conditions, it corresponds to the critical point of the system.
Scaling arises in the vicinity of this FP and the universal values of the exponents governing scaling
of different physical quantities may be found. Having this preamble in mind, we can make a brief
252
Introduction to renormalization
review of what results for the RG flows, FPs and exponents of the non-ideal models have been
found so far and how they differ from those of the “ideal” m-vector model.
6.1. m-vector magnet
Put in a formal way, the question: does the 3d m-vector magnet exhibit a critical point (a Curie
point for a ferromagnet or a Néel point for an antiferromagnet), transforms into the question: is
there a stable and reachable FP in the flow equations (cf. (4.14)–(4.16)) for the couplings of its
effective Hamiltonian? The flow equation for the coupling u of the effective Hamiltonian (5.12) can
be written in a general form of an ordinary first order differential equation:
du
d ln `
= β(u), (6.1)
with the RG flow parameter ` and β-function β(u). The parameter ` may serve to measure a
distance to the critical point: ` → 0 corresponds to T → Tc, and a specific form of the β-function
depends on a choice of the RG procedure. The data discussed in this chapter have been obtained
within the field-theoretic RG [4] approach at d = 3. Giving numbers, we shall refer to the source,
where they were obtained. However we shall try to be not too specific, mentioning several review
papers for the interested reader.
1
0
-1
1.510.50
β(u) 135
4 2
u
(a)
1
0
-1
210
β(u) 1 3 5 4
2
u
(b)
Figure 4. β-function of the 3d m = 1 model in successive perturbation theory orders ranging
from 1 to 5 as shown by the labels in the figures. Left: naive evaluation of the function (6.2).
Right: resummation of (6.2) taking into account asymptotic properties of the series. Note a
decrease of the difference between FP coordinates found in successive orders of the resummed
perturbation expansion, figure 4b.
To give an idea about the RG expansions and their treatment, we write down several first terms
of the β-function (6.1) [26]:
β(u) = −u(ε − u + 3u2(3m + 14)/(m + 8) + · · · ), (6.2)
with ε = 4 − d. For a quantitative analysis, one can develop an ε-expansion looking for solutions
of the FP equation
β(u∗) = 0 (6.3)
in a form of a series u∗ =
∑
i u(i)εi. Alternatively, one can solve the non-linear equation (6.3)
directly at d = 3 (ε = 1). However, as it is well known by now, the perturbative RG expansions
have zero radius of convergence and are asymptotic at best [4] (cf. behaviour of the function (6.2),
figure 4a). Special procedures of resummation have been elaborated to deal with them. We give
an example of how one of these procedures works transforming a divergent series, figure 4a, into
253
Yu.Holovatch
a convergent one, figure 4b. In the last figure, the function (6.2) has been resummed by the Padé-
Borel resummation. The procedure consists of several steps. First, assuming the factorial growth
of the coefficients ci in the expansion β(u) =
∑
i ciu
i (6.2), one constructs the Borel trransform of
the initial function β(u) via:
βB(u) =
∑
i
ciu
i
i!
. (6.4)
Then, the Borel transform is extrapolated by a Padé approximant [K/L](u). The last is the ratio
of two polynomials of order K and L such that its truncated Taylor expansion is equal to βB(u).
The resummed function is then calculated by an inverse Borel transform of this approximant:
β(u) =
∫ ∞
0
dt exp(−t)[K/L](ut). (6.5)
Similar techniques are currently widely used in analyzing the RG expansions [4,27]. In particular,
the numerical estimates of different physical quantities given in this chapter (and in Table 1 as
well) have been obtained using resummation techniques.
One sees the presence of two FPs in figure 4b: an unstable Gaussian FP G u∗ = 0 (the
slope of the β-function is negative, ∂β(u)/∂u|u∗ < 0) and a stable one, u∗ 6= 0. Here, the slope
of the β-function is positive: starting form any initial conditions with u > 0 the solution of the
differential equation (6.1) reaches its FP value. This FP corresponds to the critical point Tc of
the 3d Ising model. Similar behaviour of the β-function is found for other values of m, therefore
the FP with u 6= 0 for general m is called a O(m)-symmetric or Heisenberg FP H. We do not
show the procedure of calculating the critical exponents: the other RG functions being evaluated
at this FP bring about the asymptotic values of the critical exponents (in particular, those given
in Table 1). These exponents govern criticality of systems of different microscopic nature, which
can be described by the effective Hamiltonian (5.12). It is said that these systems belong to the
O(m) universality class. In the RG picture, different microscopic origin is reflected in different
initial conditions for the flow. However, the FP location and stability is defined solely by the global
features: dimensionality, symmetry, interaction type.3 Systems which share the global features
belong to the same universality class.
6.2. Cubic model
The main difference of the effective Hamiltonian of the cubic model (5.14) as compared to
that of the m-vector model (5.12) is that it contains one more coupling v of different symmetry.
Therefore, two β-functions describe the RG flow:
du
d ln `
= βu(u, v),
dv
d ln `
= βv(u, v). (6.6)
Stability of a FP is now defined by the stability matrix composed of the β-function derivatives
∂βu,v/∂{u, v}. The FP picture, which arises form an analysis of the β-functions at d = 3 [10] is
schematically shown in figure 5. Four FPs are obtained: unstable Gaussian G (u = v = 0) and
Ising I (u = 0, v 6= 0) as well as Heisenberg H (u 6= 0, v = 0) and mixed M (u 6= 0, v 6= 0). The
stability of the two last FPs depends on the value of m.
Two different regimes for the RG flows are observed. For small m < mcub
c the FP H is stable. At
this FP the system does not feel a presence of the cubic coupling, v = 0. Therefore, for m < mcub
c
the cubic model belongs to the O(m) universality class: its exponents coincide with the exponents
of the m-vector model, table 1. However, with an increase of m the FP M approaches H and at
m = mcub
c both FPs coincide: a crossover to the new regime occurs. For m > mcub
c the FP M
becomes stable and governs the critical properties of the cubic model in the new universality class.
The marginal dimension value is slightly less than three: mcub
c = 2.862(5) [28]. From this estimate
3In particular, here we consider systems with the short-range interaction.
254
Introduction to renormalization
v
I
G H
M
u
(a)
v
I
G H
M
u
(b)
Figure 5. FPs and RG flows of the cubic model. Unstable FPs are shown by discs, stable FPs
are shown by squares. (a): m < mcub
c , the Heisenberg FP H is stable. (b): m > mcub
c , the mixed
FP M is stable. The dotted line in figure 5a shows a separatrix: the RG flows which start below
this line do not reach H. Note that for m > mcub
c the u-axis is a separatrix: all flows that start
from the initial conditions with v < 0 do not reach any FP (run-away solutions).
it follows in particular, that a cubic Heisenberg (m = 3) magnet does not belong to the O(3)
universality class. Its critical exponent being estimated in the fixed point M read [29]:
γ = 1.390(12), ν = 0.706(6), η = 0.0333(26), β = 0.364(15), α = −0.118(18). (6.7)
Numerically, these values are close to their counterparts for the m-vector model (cf. Table 1).
However the principal difference arises from the above analysis: as one can easily check solving
the system of differential equations (6.6) for m > mcub
c the RG flows with v < 0 cannot reach the
stable FP (an abscissa v = 0 serves as a separatrix for the flows). As it follows from sections 3.2, 5.2
negative v corresponds to ordering along diagonals of m-dimensional hypercube. Therefore, the RG
analysis results in a statement that such ordering cannot occur via a 2nd order phase transition:
ferromagnetic crystals with three easy axes should undergo a 1st order phase transition. It is
worth noting here, that whereas the presence of stable and reachable FP brings about the 2nd
order phase transition, its absence signalling only that the 2nd order phase transition does not
occur. The nature of the low-temperature phase and the scenario of how it is attained remains to
be checked by other methods.
6.3. Random-site dilution
Formally, an analysis of the weakly diluted quenched m-vector model resembles those we dis-
cussed in the former subsection 6.2. Indeed, both effective Hamiltonians (5.14), (5.25) contain two
couplings of different symmetry and give rise to the already familiar FP picture. Moreover, the
effective Hamiltonian (5.25) at m = 1 coincides with (5.14).4 However, our goal is to analyze it in
the replica limit n = 0. Note, that now the physically meaningful values of couplings are u > 0,
v < 0, see section 5.3. The FP picture and the RG flows at d = 3 are shown in the figure 6 [13].
Besides the familiar FPs G, H, and M a polymer FP P is present. It is stable and corresponds
to the O(m = 0) universality class, however it is never reached from the initial conditions u > 0,
v < 0.
Again, the new marginal dimension mdil
c governs the crossover between the new and the O(m)
universality classes: FP M is stable for m < mdil
c . A search for the value of mdil
c shows, that at
m = mdil
c the heat capacity critical exponent of an undiluted system changes its sign: α(mdil
c ) = 0.
In this way one recovers the Harris criterion [17], section 3.3.1, translated into the RG “language”.
4One can check it by further substitution n → m, {u0, v0} → {v0, u0}.
255
Yu.Holovatch
v
P
G H
M
u
(a)
v
P
G H
M
1
2
3
u
(b)
Figure 6. FPs and RG flows of the weakly diluted quenched m-vector model. Unstable FPs are
shown by discs, stable FPs are shown by squares. (a): m < mdil
c , the mixed FP M is stable. (b):
m > mdil
c , the Heisenberg FP H is stable. Flows 1, 2, 3 of the figure 6b are further treated in
figure 7b.
The numerical value of mdil
c being slightly less than two, mdil
c = 1.912(4) [30], only the Ising
model (m = 1) changes its exponents upon dilution. Indeed, the numerical values of the exponents
read [31]:
γ = 1.330(17), ν = 0.678(10), η = 0.030(3), β = 0.349(5), α = −0.034(30) (6.8)
and differ essentially from those of the 3d Ising model (table 1).
However, both in the experiments and in the MC simulations one deals with the system not
yet in an asymptotic region, where exponents do not attain their FP values and the effective
exponents are found. Being non-universal, they can be calculated in the RG treatment as functions
of the flow-dependent couplings. In figure 7, we show an effective critical exponent γeff measured
recently for the ac susceptibility of the a-Fe86Mn4Zr10 amorphous alloy [32] and compare them with
the theoretical RG calculations of the effective exponents of the weakly diluted quenched m = 3
model. The exponent was calculated along different RG flows labeled by numbers in figure 6b. This
calculation serves an example how the non-asymptotic effects may be taken into account in the RG
analysis. Although direct correspondence between the temperature distance to the critical point τ
and the RG flow parameter ` is problematic, the RG serves as a useful tool of accompanying the
studies of effective critical behaviour.
6.4. Random anisotropy
On this example we shall show how the RG predicts two different phenomena occuring in the
random anisotropy magnets. It appears that the type of local random axis distribution crucially
effects an origin of the low-temperature phase in random anisotropy systems [19]. First, we consider
the results obtained for an isotropic distribution (3.7), which leads to the effective Hamiltonian
(5.29) [21]. It contains three couplings, u, v, w, hence three β-functions define the RG flows. Solving
the FP equations one arrives at the FP picture shown in figure 8a [21,20].
Let us recall (cf. section 5.4), that physically meaningful values of couplings are u > 0,
v > 0, w < 0. Therefore only the FPs located in this region are shown in the figure. However,
there is another condition found for the ratio of couplings: w0/u0 = −m. The region of typical
initial conditions to study the RG flow is shown in figure 8 by a cross. The only stable FP found,
a polymer FP P is not reachable from the initial conditions. The run-away solutions of the RG
equations bring about an absence of a 2nd order phase transition.
A different picture is obtained for a cubic random axis distribution (3.8) [20]. Here, the effective
Hamiltonian (5.30) contains four couplings of different symmetry, u, v, w, y. The physical initial
256
Introduction to renormalization
γeff
τ
(a)
γeff
ln `
1
2
3
(b)
Figure 7. Effective critical exponent γeff for the magnetic susceptibility of the weakly diluted
quenched Heisenberg magnet. (a): as a function of the distance to the Curie point τ = (T−Tc)/Tc
in experimental measurements for a-Fe86Mn4Zr10 amorphous alloy [32]. (b): as a function of the
RG flow parameter in theoretical RG calculations for the diluted m = 3 model [15]. Different
curves correspond to different amount of disorder. Note that in the asymptotics (τ → 0 or ` → 0)
the exponent attains its universal value.
values for the couplings lay in the region (section 5.4): u > 0, v > 0, w < 0, w/u = −m. A typical
FP picture is shown in figure 8b for v = 0. Similar to the former case of isotropic random axis
distribution, the stable FP P cannot be reached. However, one more stable FP M is present. It is
reachable for the RG flow that starts from the initial conditions marked by a cross in the figure.
This FP is a FP of the random-site Ising model (subsection 6.3) for any value of m. It means that
the ferromagnetic 2nd order phase transition in the m-vector magnet with the cubic random axis
distribution belongs to the universality class of the random-site Ising model and is governed in
asymptotics by the exponents (6.8).
6.5. Stacked triangular antiferromagnet
Again, as in the former subsections 6.1–6.4, the RG answer about a possibility of a 2nd order
phase transition in the 3d stacked triangular antiferromagnet would be a presence of a stable
accessible FP for the couplings u, v of the effective Hamiltonian (5.32). The model possesses a
rather complicated FP structure, sketched in figure 9.
FP picture changes with m and one finds three marginal dimensions mchir
i that govern its
topology. For large m > mchir
3 the stable chiral FP C+ is present and it can be reached from the
initial values of the couplings u, v > 0. Therefore, the phase transition to the non-collinear chiral
state is of the second order. At m = mchir
3 , C+ merges with the anti-chiral FP C− and disappears,
only one unstable Heisenberg FP H is found for mchir
3 > m > mchir
2 . With further decrease of m,
two more FPs appear at m = mchir
2 , S− and S+. The last one is stable, however it describes the
sinusoidal phase, which occurs for v < 0. Finally, at m = mchir
1 this FP merges with H and looses
its stability with further increase of m. It follows from this analysis, that the 2nd order phase
transition into non-collinear phase can occur only in the magnets with m > mchir
3 .
In principle, the above described FP picture has been known since the first RG studies of
the problem [22]. However, reliable numerical estimates for marginal dimensions mchir
i have been
obtained only recently. The question of interest is, what is the order of the phase transition at m =
2; 3 when the model has its physical realization? The estimates: mchir
3 = 6.23(21), mchir
2 = 1.99(4),
mchir
1 = 1.43(2) [33] clearly rule out the possibility of a 2nd order phase transition for m = 2; 3:
the FP picture is shown in the second figure of figure 9 and no accessible FP is found.
Note, however, certain controversy in the RG studies performed so far: whereas the above FP
picture is confirmed by the perturbative RG expansions and the non-perturbative RG approach
257
Yu.Holovatch
(a) (b)
Figure 8. Fixed points of the random anisotropy model. The filled boxes show the stable fixed
points, the cross denotes typical initial values of couplings. (a): isotropic local random axis
distribution. The stable FP P cannot be reached by the RG flow which starts from the region
shown by a cross in the figure. (b): cubic local random axis distribution. The stable FP P cannot
be reached, but the random Ising FP M is stable and reachable for the RG flow.
[23,33], an analysis of the FP equations by direct solution of the non-linear FP equations for the
resummed β-functions brings about a presence of the stable reachable FP for m = 2; 3 [34]. The
last is associated with the critical point of the 2nd order phase transition.
7. Conclusions and outlook
How do the changes in structure effect the critical behaviour of the matter? We tried to give
an answer to this question taking a 3d m-vector model as an ideal reference system and showing
what will happen to it under an effect of different non-idealities. The examples considered include
anisotropy, structural disorder, frustrations: the features one often encounters dealing with realis-
tic condensed matter objects. The response of a system to such non-idealities appears to be very
different, ranging from an insensitivity (the cubic m = 2 magnet remains in the O(2) universality
class), through softening (the heat capacity of the random site m = 1 magnet does not diverge)
to disappearance of a 2nd order phase transition (isotropically distributed local random axis de-
stroys long-range order). These various asymptotic features are accompanied by a complicated
non-asymptotic effective critical behaviour.
It is astonishing that all this bunch of phenomena can be explained and accurately described
within one theoretical framework, the RG approach. An application of basic RG notions of flows,
fixed points and their accessibility, marginal dimensions and crossovers supported by an elaborate
machinery to perform and analyze the RG transformation resulted in a coherent picture of phe-
nomena in the vicinity of a critical point. A lot remains to be cleared up in this picture. Maybe a
participant of the school or a reader of these lectures will decide to make his or her contribution?
Good luck!
Acknowledgements
I am grateful to Bertrand Berche, Arnaldo Donoso, and Ricardo Paredes for the invitation
to lecture at the Spring school on Foundations of statistical and mesoscopic physics (Mochima,
Venezuela, June 20th–24th 2005) and to all participants of the school for the wonderful atmo-
sphere created there. Bertrand Berche is further acknowledged for his encouragement, advice during
preparation of the manuscript and, last but not least, for his stories about the Cagniard de la Tour
258
Introduction to renormalization
m > m
chir
3
v
u
G H
C−
C+
v
u
G H
m
chir
3
> m > m
chir
2
v
G H
S+
S−
u
m
chir
2 > m > m
chir
1
v
G H
S+
S−
u
m < m
chir
1
Figure 9. FPs and RG flows of the stacked triangular antiferromagnet model. Unstable FPs are
shown by discs, stable FPs are shown by squares. Three marginal dimensions mchir
1 , mchir
2 , mchir
3
govern the FP picture.
state, the critical state discovered as early as in 1822 [1]! I thank my colleagues Viktoria Blavats’ka,
Bertrand Delamotte, Maxym Dudka, Christian von Ferber, Reinhard Folk, Dmytro Ivaneiko, Taras
Yavors’kii – some of the results mentioned in the last part of these lectures are due to our common
work.
This work was supported by Austrian Fonds zur Förderung der wissenschaftlichen Forschung
under Project No. P16574.
References
1. A historical overview of the experimental and theoretical studies of critical phenomena may be found
in J.S. Rowlinson introductary essay to the English edition of J.D. van der Waals Leiden thesis of 1873:
J.D. van der Waals: On the Continuity of the Gaseous and Liquid States. Edited with an Introductory
Essay by J.S. Rowlinson. North-Holland Physics Publishing, Amsterdam, 1988; and in: C. Domb. The
Critical Point. Taylor & Francis, London, 1996.
2. J.E. Verschaffelt. Koninklijke Akademie van Wetenschappen, Proc. Sect. Scienc., 1990, II, 588. English
version: http://www.knaw.nl/waals/pdf/Verschaffelt01.pdf
259
Yu.Holovatch
3. Wilson K.G., Kogut J., Phys. Rep., 1974, 12, 75.
4. The field theoretical RG in its applications to the phase transition theory is explained in: E. Brézin,
J. C. Le Guillou, and J. Zinn-Justin, in: Phase Transitions and Critical Phenomena, edited by C. Domb
and M. S. Green, Vol. 6, Academic Press, London, 1976; Amit D.J. Field Theory, the Renormalization
Group, and Critical Phenomena. World Scientific, Singapore, 1989; Zinn-Justin J. Quantum Field
Theory and Critical Phenomena. Oxford University Press, Oxford, 1996.
5. Ising E., Z. Phys., 1925, 31, 255.
6. Stanley H.E., Phys. Rev. Lett., 1968, 20, 589.
7. Mermin N.D., Wagner H., Phys. Rev. Lett., 1966, 17, 1133; Erratum: ibid., 1966, 17, 1307;
Hohenberg P.C., Phys. Rev., 1967, 158, 383.
8. Onsager L., Phys. Rev., 1944, 65, 117.
9. Guida R., Zinn-Justin J., J. Phys. A, 1998, 31, 8103.
10. Aharony A. Phase Transitions and Critical Phenomena, edited by C. Domb and M. S. Green, Vol.6.
Academic Press, London, 1976.
11. Brout R., Phys. Rev., 1959, 115, 824.
12. Belanger D.P., Young A.P., J. Magn. Magn. Mater., 1991, 100, 272.
13. A review of the random 3d Ising model critical behaviour may be found in: Folk R., Holovatch Yu.,
Yavors’kii T., Physics – Uspekhi, 2003, 46, 169 [Uspekhi Fizicheskikh Nauk, 173, 175] and in [14].
14. A review on the critical behaviour of systems with complex effective Hamiltonians: Pelissetto A.,
Vicari E., Phys. Rep., 2002, 368, 549.
15. Dudka M., Folk R., Holovatch Yu., Ivaneiko D., J. Magn. Magn. Mater., 2003, 256, 243.
16. Yoon J., Chan M.H.W., Phys. Rev. Lett., 1997, 78, 4801.
17. Harris A.B., J. Phys. C, 1974, 7, 1671.
18. Harris R., Plischke M., Zuckermann M.J., Phys. Rev. Lett., 1973, 31, 160.
19. A review of early work on random anisotropy magnets may be found in: Cochrane R.W., Harris R.,
Zuckermann M.J., Phys. Rep., 1978, 48, 1. Recent experimental, numerical, and theoretical studies
are reviewed in [20].
20. Dudka M., Folk R., Holovatch Yu., Journ. Mag. Mag. Mat., 2005, 294, 305.
21. Aharony A., Phys. Rev. B, 1975, 12 1038.
22. Kawamura H., Phys. Rev. B, 1988, 38 4916.
23. A review on frustrated magnets with an emphasis on the results obtained within the non-perturbative
RG approach: Delamotte B., Mouhanna D., Tissier M., Phys. Rev. B, 2004, 69, 134413.
24. Fisher M.E., in: Critical Phenomena. Proceedings of the Summer School, Stellenbosch, South Africa.
Ed. by F. J. W. Hahne. Lecture Notes in Physics, 186, Springer Verlag, Berlin, 1983; D. R. Nelson,
Fisher M.E., Ann. Phys. (N.Y.), 1975, 91, 226.
25. Emery V.J., Phys. Rev. B, 1975, 11, 239; Edwards S.F., Anderson P.W., J. Phys. F, 1975, 5, 965.
26. Kleinert H., Neu J., Schulte-Frohlinde V., Chetyrkin K.G., Larin S.A., Phys. Lett. B, 1991, 272, 39;
Erratum: ibid., 1993, 319 545.
27. A review on application of resummation technique to divergent RG expansions of disordered models:
Holovatch Yu., Blavats’ka V., Dudka M., von Ferber C., Folk R., Yavors’kii T., Int. J. Mod. Phys. B,
2002, 16, 4027.
28. Folk R., Holovatch Yu., Yavors’kii T., Phys. Rev. B, 2000, 62, 12195.
29. Carmona J.M., Pelissetto A., Vicari E., Phys. Rev. B, 2000, 61, 15136.
30. Holovatch Yu., Dudka M., Yavors’kii T., J. Phys. Stud., 2001, 5, 233.
31. Pelissetto A., Vicari E., Phys. Rev. B, 2000, 62, 6393.
32. Perumal A., Srivinas V., Rao V.V., Dunlap R.A., Phys. Rev. Lett., 2003, 91, 137202.
33. Holovatch Yu., Ivaneiko D., Delamotte B., J. Phys. A, 2004, 37, 3569.
34. Pelissetto A., Rossi P., Vicari E., Phys. Rev. B, 2001, 63, 140414.
Questions and answers
Q (Alexander López): How do the signs of the couplings u and v in the effective Hamiltonian
(5.14) determine a type of the low-temperature ordering?
A Neglecting fluctuations (taking a function φ(r) to be just a variable φ) you can think about
the effective Hamiltonian as of the Landau free energy. Now, let the reference system display
a 2nd order phase transition. This means that u > 0. Minimizing Landau free energy and
looking for the spontaneous magnetization one finds at T < Tc two different non-trivial
260
Introduction to renormalization
solutions: ~φ = (φ/
√
m, . . . , φ/
√
m), it exists at v < 0, and ~φ = (φ, 0, . . . , 0) at v > 0. They
correspond to two types of ordering: along the diagonals or along the edges of a m-dimensional
hypercube. For m = 3, these are directions [111] and [100] correspondingly.
Q (Bertrand Berche): What are experimental realizations of the cubic model?
A I have already mentioned ferromagnetic crystals. Besides, at m = 3 the model describes a
ferroelectric phase transition which occurs in SrTiO3 at 105 K ( Cowley R.A., Bruce A.D.,
J. Phys. C, 1973, 6, L191). Moreover, since the model provides an example of a system with
an arbitrary weak first-order phase transition, it is also used as a testing ground to describe
an elecroweak transition in the early Universe (P. Arnold, S Sharpe, L. Yaffe, Y. Zhang, Phys.
Rev. Lett., 1997, 78, 2062).
Q (Carlos Vásquez): In 1983, Weinrib and Halperin proposed a model to describe an effect of
extended (correlated) disorder on magnetic 2nd order phase transition. There, the impurity-
impurity correlation function decays for large separations as g(r) ∼ r−a. Currently, there exist
two different predictions for the critical exponents of such a model. What are the methods
used to obtain them?
A Indeed, the original result of Weinrib and Halperin ( A. Weinrib, Halperin B.I., Phys. Rev. B,
1983, 27, 417) was obtained in the first order of the expansion in ε = 4−d, δ = 4−a. Recent
two-loop estimates (V. V. Prudnikov, Prudnikov P.V., Fedorenko A.A., J. Phys. A, 1999, 32,
L399) are due to the fixed d, a RG technique. Qualitative answer of both approaches is that
for a < d the disorder is relevant if the correlation length critical exponent of the system
without defects obeys ν < 2/a.
Q (Bertrand Berche): Could you comment on logarithmic corrections to the scaling laws? Is the
critical behaviour at marginal dimensions you were speaking about (e.g. at mdil
c ' 1.91 for
the 3d diluted m-vector magnet) governed by such corrections?
A Logarithmic corrections arise at upper critical dimension dup: that is, at the space dimension,
above which the mean-field theory holds5. For the models I was speaking about, the upper
critical dimension is four. Indeed, the divergence of an isothermal susceptibility of the m-
vector model at d = 4 is governed by the mean field exponent γ = 1. However, the power
law singularity is accompanied by a logarithmic one:
χ−1 ∼ |τ |−1ln |τ |−(m+2)/(m+8)
, (7.1)
as first derived in: Larkin A.I., Khmelnitskii D.E., JETP, 1969, 29, 1123. In the RG scheme,
one can see the origin of such corrections solving the flow equation (6.1) at d = 4 (ε = 0):
du
d ln `
= −u2, (7.2)
here the right-hand side is the leading term of the β-function (6.2). The solution of equa-
tion (7.2)
u =
1
| ln `| + const, ` → 0 (7.3)
being substituted into an appropriate expression for the susceptibility leads to the above
dependence (7.1). Now, let us return to the marginal dimensions mc we were discussing in
these lectures. They are field dimensions and we estimated them at space dimension d = 3.
As far as the space dimension was lower than dup = 4, the logarithmic corrections do not
appear. Returning back to the example given by formulas (7.2), (7.3): even at m = mc, the
first power of couplings is present in the right-hand side of functions (7.2) for d = 3 and the
solutions will rather behave as u ∼ `const.
5Note added in proof: We do not discuss here the logarithmic corrections appearing at low dimensions as those
in 2d diluted Ising model or 2d q = 4 Potts model. For a list of systems where logarithmic corrections appear and
for the scaling relations between them see: Kenna R., Johnston D.A., Janke W., Phys. Rev. Lett., 2006, 96, 115701.
261
Yu.Holovatch
Q (Dragi Karevski): Do you know other citeria, similar to the Harris one, which predict changes
in the critical behaviour caused by different types of disorder? For example, what happens
when disorder is coupled to the order parameter?
A Indeed, Harris criterion concerns the systems, where disorder is coupled to the energy density
(look for example at the Hamiltonian (3.3), where random variables cR are coupled to the
product of spins). It states that critical exponents of a disordered system do not change, if
the heat capacity of the pure system does not diverge. Later, a statement that the correlation
length critical exponent of d-dimensional systems with such type of disorder should obey an
inequality ν > 2/d was proven ( Chayes J.T., L. Chayes, Fisher D.S., T. Spenser, Phys.
Rev. Lett., 1986, 57, 2999). For the extended, long-range correlated disorder the generalized
Harris criterion holds (see the above mentioned paper of Weinrib and Halperin as well as D.
Boyanovsky, Cardy J.L., Phys. Rev. B, 1983, 27, 6971).
Y. Imry and S.-k Ma have shown (Phys. Rev. Lett., 1975, 35, 1399) that even arbitrary weak
disorder coupled to the order parameter of continuous symmetry (i.e. for m > 2 vector model)
destroys ferromagnetism at d < 4. For the random-field Ising model (m = 1) the lower critical
dimension is d = 2. As I already have mentioned in the lectures, the isotropically distributed
random axis destroys ferromagnetism at d < 4. One of the ways of showing this is to exploit
the arguments similar to those of Imry and Ma for the random-field systems (Pelcovits R.A.,
Pytte E., Rudnick J., Phys. Rev. Lett., 1978, 40, 476).
Last but not least, let me mention the Luck criterion for the connectivity disorder, as ex-
plained in the lecture by Wolfhard Janke.
����� �� �� �
� ���
�
�
.�������1,2,3
1 I������� �i���� ���� ���!"��# ���� $ %&% '�(")��, 79011 *+!i!, !�,. -!.�/i/+��0�, 1, '�(")�"
2 I������� � �( ��1��) �i����, '��! (��� � 2�0"�" 3 4, (" $i��" *i�/, *i�/ 4040,
&!��(i5
3 *+!i!�+��6 �"/i��",+��6 ��i! (��� � i$. I.7("��", *+!i! 79005, '�(")�"
89:;<=>? 13 @A9?B? 2006 :., C ?D9=9?E>?<F C;B@GHi – 17 ICi9>G 2006 :.
J /�# , �/i5# 5 �K0�!�(LL ���K,�!���i �(���1��) 4�! �i��� “� i� ",+��#” ���� $ i � , 5� )# $�-M�" 4�5����� �" ��4�$�0�L$ ���� ( ��($",i�"/i6��) 0(�4�. N��0,5�"L�+�5 ���� $� � ����(i���L"�i���(�4i.L, ��(����(��$ K �,"��$, �(���("/i5$�. O �"4(�!"�M�L ����!�i i� ) 4 ( ��($�!"��5
i 4��"��L, 5� !��� �,�M"�+ �,5 4�5�� ��5 ��4�!�# (�� �(���1����i ! �"��# ���� $"#: 4�$’5�P ��5�"��!�0� 4 ( #���, �$i�� �,"�� ��i! (�",+����i, ��,"���) � ���!��) �(���1��) 4�! �i���.
QRSTUVi WRUVX: YZ[\[]^_ `abcd i^Y_, `cZc^aZefb_^^g, \caZc\[Ya-`ahiab_ Zc^aZe_hij_kil^_ mZf`_
PACS: 05.50.+q, 05.70.Jk, 64.60.Ak
262
|