On Affine Fusion and the Phase Model
A brief review is given of the integrable realization of affine fusion discovered recently by Korff and Stroppel. They showed that the affine fusion of the su(n) Wess-Zumino-Novikov-Witten (WZNW) conformal field theories appears in a simple integrable system known as the phase model. The Yang-Baxter...
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| citation_txt | On Affine Fusion and the Phase Model / M.A. Walton // Symmetry, Integrability and Geometry: Methods and Applications. — 2012. — Т. 8. — Бібліогр.: 19 назв. — англ. |
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| description | A brief review is given of the integrable realization of affine fusion discovered recently by Korff and Stroppel. They showed that the affine fusion of the su(n) Wess-Zumino-Novikov-Witten (WZNW) conformal field theories appears in a simple integrable system known as the phase model. The Yang-Baxter equation leads to the construction of commuting operators as Schur polynomials, with noncommuting hopping operators as arguments. The algebraic Bethe ansatz diagonalizes them, revealing a connection to the modular S matrix and fusion of the su(n) WZNW model. The noncommutative Schur polynomials play roles similar to those of the primary field operators in the corresponding WZNW model. In particular, their 3-point functions are the su(n) fusion multiplicities. We show here how the new phase model realization of affine fusion makes obvious the existence of threshold levels, and how it accommodates higher-genus fusion.
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Symmetry, Integrability and Geometry: Methods and Applications SIGMA 8 (2012), 086, 13 pages
On Affine Fusion and the Phase Model?
Mark A. WALTON
Department of Physics and Astronomy, University of Lethbridge,
Lethbridge, Alberta, T1K 3M4, Canada
E-mail: walton@uleth.ca
Received August 01, 2012, in final form November 08, 2012; Published online November 15, 2012
http://dx.doi.org/10.3842/SIGMA.2012.086
Abstract. A brief review is given of the integrable realization of affine fusion discovered
recently by Korff and Stroppel. They showed that the affine fusion of the su(n) Wess–
Zumino–Novikov–Witten (WZNW) conformal field theories appears in a simple integrable
system known as the phase model. The Yang–Baxter equation leads to the construction
of commuting operators as Schur polynomials, with noncommuting hopping operators as
arguments. The algebraic Bethe ansatz diagonalizes them, revealing a connection to the
modular S matrix and fusion of the su(n) WZNW model. The noncommutative Schur
polynomials play roles similar to those of the primary field operators in the corresponding
WZNW model. In particular, their 3-point functions are the su(n) fusion multiplicities. We
show here how the new phase model realization of affine fusion makes obvious the existence
of threshold levels, and how it accommodates higher-genus fusion.
Key words: affine fusion; phase model; integrable system; conformal field theory; noncom-
mutative Schur polynomials; threshold level; higher-genus Verlinde dimensions
2010 Mathematics Subject Classification: 81T40; 81R10; 81R12; 17B37; 17B81; 05E05
1 Introduction
Affine fusion is a natural generalization of the tensor product of representations of simple Lie
algebras. It is a simple truncation thereof controlled by a non-negative integer, the level. As such,
it is a basic mathematical object, found in many different mathematical and physical contexts.
The physical context preferred by this author is provided by conformal field theory, and the so-
called Wess–Zumino–Novikov–Witten (WZNW) models (see [4], for example). WZNW models
realize at a fixed non-negative integer level k a non-twisted affine Kac–Moody algebra g(1) based
on a simple Lie algebra g, or gk for short. Their primary fields furnish representations of gk and
their operator products are governed by the corresponding affine fusion algebra.
Recently, Korff and Stroppel [16] found a much simpler physical realization of affine fusion,
for the su(n)k case. The phase model [2] is an integrable multi-particle model whose integrals of
motion may be diagonalized by the algebraic Bethe ansatz [13, 17]. Its integrability is not only
crucial to its realization of su(n)k fusion, but also explains certain properties. The integrable, or
phase-model realization of affine fusion raises hope that a better understanding of affine fusion
and its physical contexts will result from its study.
This paper is meant to be a gentle, non-rigorous introduction to the phase-model realization
of affine fusion. We hope that others will share our interest in the topic and the mathematical
tools involved, and perhaps help develop them further. Other reviews can be found in [14, 15].
Sections 2–4 constitute the introductory review. Section 5 contains some new results: thre-
shold levels (and threshold multiplicities and polynomials) and higher-genus Verlinde dimensions
are both treated in the phase model there for the first time. Section 6 is a short conclusion.
?This paper is a contribution to the Special Issue “Superintegrability, Exact Solvability, and Special Functions”.
The full collection is available at http://www.emis.de/journals/SIGMA/SESSF2012.html
mailto:walton@uleth.ca
http://dx.doi.org/10.3842/SIGMA.2012.086
http://www.emis.de/journals/SIGMA/SESSF2012.html
2 M.A. Walton
2 Phase model: Hilbert space and operator algebras
The set of highest weights λ of integrable highest-weight representations L(λ) of su(n) is
P+ :=
{
(λ1, λ2, . . . , λn−1) :=
n−1∑
a=1
λaΛ
a |λa ∈ Z≥0
}
,
where Λa is the a-th fundamental weight. Identifying this su(n) as the horizontal subalgebra of
the affine Kac–Moody algebra su(n)k at level k,
P k+ :=
{
λ = [λ1, . . . , λn−1, λn] :=
n∑
a=1
λaΛ
a |λa ∈ Z≥0,
n∑
a=1
λa = k
}
is the set of affine highest weights at level k.
The phase model has a Hilbert space H with basis labeled by affine highest weights: |λ〉 =
|λ1, . . . , λn−1, λn〉. The Dynkin labels are interpreted as the numbers of particles at n sites on
a circle, corresponding to the nodes of the affine Dynkin diagram. If Na denotes the number
operator for site a, then the level k is the total number of particles
Na|λ〉 = λa|λ〉 ⇒ N |λ〉 = k|λ〉, N :=
n∑
a=1
Na.
The basis of states is orthonormal: 〈λ|µ〉 = δλ,µ. Notice that this means states of different levels
(numbers of particles) are orthogonal.
Define operators ϕ†i and ϕi that create and annihilate (respectively) particles at site i:
ϕ†i |λ〉 = | . . . , λi−1, λi + 1, λi+1, . . .〉,
ϕi|λ〉 =
{
| . . . , λi−1, λi − 1, λi+1, . . .〉, λi ≥ 1,
0, λi = 0.
In the phase model, these operators obey the so-called phase algebra [2, 16], generated
by ϕ†i , ϕi and the number operators Ni, for i ∈ {1, . . . , n}, with relations
[ϕi, ϕj ] = [ϕ†i , ϕ
†
j ] = [Ni, Nj ] = 0, [Ni, ϕ
†
j ] = δi,jϕ
†
i , [Ni, ϕj ] = −δi,jϕi,
Ni(1− ϕ†iϕi) = 0 = (1− ϕ†iϕi)Ni, [ϕi, ϕ
†
j ] = 0 if i 6= j, but ϕiϕ
†
i = 1. (1)
Notice that the commutator of ϕi and ϕ†i does not appear in the defining relations of this
algebra. That’s because the phase model is the crystal limit of the q-boson hopping model, as
made precise in [14], so that a q-commutator reduces at q = 0 to the last relation of (1). The
operator πi := (1− ϕ†iϕi) projects to states with no particles at site i (so π2
i = πi).
As already mentioned, the level becomes the total particle number here. Therefore, to realize
the fusion of a WZNW model, which has a fixed level k, we must restrict to a fixed total particle
number. Hopping operators
ai := ϕ†iϕi−1, i ∈ {1, . . . , n} (2)
are then important1. Here the indices are defined mod n, so that a1 = ϕ†1ϕn.2 The action of ai
is
ai|λ〉 =
{
| . . . , λi−2, λi−1 − 1, λi + 1, λi+2, . . .〉 = |λ− Λi−1 + Λi〉, λi−1 ≥ 1,
0, λi−1 = 0.
(3)
1When their action is non-trivial, the operators ϕ†i−1ϕi = a†i hop particles in the opposite direction around
the sites of the affine su(n) Dynkin diagram. We will focus on the ai.
2For simplicity, we put the “magnetic flux parameter” z of [16] to 1.
On Affine Fusion and the Phase Model 3
The algebra of the hopping operators A = 〈a1, a2, . . . , an〉 is defined by the relations
A : [ai, aj ] = 0 if i 6= j ± 1 mod n,
aia
2
j = ajaiaj & a2
i aj = aiajai if i = j + 1 mod n, (4)
easily verified from the phase algebra (1), in view of (2). This algebra A is called the affine
local plactic algebra in [16]. The first line of (4) is the locality condition. Plactic algebras were
first defined by Lascoux and Schützenberger, and named by them because of their relation to
tableaux (see [7]).
For n = 3, the relations are
a2
3a2 = a3a2a3, a3a
2
2 = a2a3a2, a2
2a1 = a2a1a2, a2a
2
1 = a1a2a1,
a2
1a3 = a1a3a1, a1a
2
3 = a3a1a3. (5)
Notice that there are no locality relations for this case – each node on the Dynkin diagram is
a nearest neighbour of the other 2.
When the indices of the relations defining A are not identified mod n, the algebra becomes
the local plactic algebra Ā = 〈a1, a2, . . . , an〉:3
Ā : [ai, aj ] = 0 if i 6= j ± 1,
ai+1a
2
i = aiai+1ai & a2
i+1ai = ai+1aiai+1, 1 ≤ i ≤ n− 1. (6)
This algebra is relevant to Young tableaux and the Littlewood–Richardson algorithm that com-
putes su(n) tensor product decompositions.
The algebra Ā can also be realized in terms of creation operators ϕ†i and annihilation opera-
tors ϕi obeying a phase algebra. More sites are needed, so that i ∈ {0, 1, 2, . . . , n}, but i = 0
and i = n are not identified. Then the generators are again constructed as ai = ϕ†iϕi−1, for
i ∈ {1, 2, . . . , n}. Identifying ϕ†0 and ϕ0 with ϕ†n and ϕn (respectively) then transforms the
construction for Ā to that for A.
The relations defining Ā for n = 3 are the same as (5), except that the third line is replaced
by [a1, a3] = 0. That is,
Ā ⇒ A : [a1, an] = 0 7→ [a1, a1an] = [a1an, an] = 0 (7)
summarizes the difference between the affine local plactic algebra and local plactic algebra for
n = 3. Looking at (4) and (6), we see that (7) applies for all n ≥ 3: in A a1 and an do not
commute, but the product a1an commutes with both a1 and an.
To see that plactic algebras are connected to Young tableaux, notice that the hopping operator
ai is associated with the weight Λi − Λi−1. These affine weights have horizontal parts equal to
the weights of the basic su(n) irreducible representation L(Λ1). The horizontal weight for ai is
the weight of the Young tableau i that labels a vector of L(Λ1) [7].
Now the states (vectors) of an irreducible su(n) representation of highest weight µ are in
1-1 correspondence with Young tableaux of shape µ and entries in {1, 2, . . . , n}. Such a Young
tableau is built starting with a Young diagram of shape µ, i.e. one with µ1 columns of height 1,
to the right of µ2 columns of height 2, etc., up to µn−1 columns of height n− 1. Since columns
of height n correspond to the trivial representation in su(n), they may be omitted.
The Young tableaux (also known as semi-standard tableaux) are obtained by filling the
Young diagram with entries from 1 to n, such that they increase going down columns, and do
3By abuse of notation, we use the same symbols for the generators of A and Ā. It is important to note that Ā
differs from the local plactic algebra Plfin defined in [16].
4 M.A. Walton
not decrease going across rows [7]. As an example, we display the Young tableaux for the adjoint
representation of su(3), of highest weight Λ1 + Λ2:
1 2
2
1 1
2
2 2
3
1 3
2
1 2
3
1 1
3
2 3
3
1 3
3
(8)
The arrangement is meant to remind the reader of the corresponding weight diagram. The
weights of su(n) Young tableaux are determined by their entries: if there are #i occurrences
of i , i = 1, . . . , n, its weight is
∑
i #i(Λ
i − Λi−1).
One version of the Littlewood–Richardson rule calculates the decomposition of the tensor-
product L(λ) ⊗ L(µ) as follows. Take the Young tableaux of shape λ and add them to the
Young diagram of shape µ, column-by-column, from right to left, to obtain “mixed tableaux”.
When adding each column, simply adjoin each box i to the i-th row of the mixed tableau. If
after adding a column, the mixed tableau has an invalid shape, there is no contribution from
the original Young tableau. If, on the other hand, the final mixed tableau survives, its shape ν
indicates the appearance of L(ν) in the desired decomposition.
For example, suppose λ = Λ1 + Λ2 and µ = Λ1, so that the appropriate Young diagram is
and the relevant Young tableaux are those of (8). Adding the rightmost column of 1 3
2
to
yields:
1 3
2
+ ⇒
3
a mixed tableau with an invalid shape, so this Young tableau produces no contribution to the
tensor-product decomposition. On the other hand, we find the sequence
1 2
3
+ ⇒
2
, 1
2
3
so the Young tableau 1 2
3
reveals a representation L(Λ1) in the decomposition.
Notice that adding a single box i to a Young diagram or mixed tableau of shape λ produces
a mixed tableaux of shape λ+ Λi − Λi−1 or vanishes, in precise correspondence with (3).
As an example of the Littlewood–Richardson rule, the Young tableaux of (8) may be added
to the Young diagram to verify the su(n) tensor product decomposition
L
(
Λ1 + Λ2
)⊗2
↪→ L(0)⊕ 2L
(
Λ1 + Λ2
)
⊕ L
(
3Λ1
)
⊕ L
(
3Λ2
)
⊕ L
(
2Λ1 + 2Λ2
)
. (9)
To understand the connection of the plactic algebra with the Littlewood–Richardson rule
and Young tableaux, we must introduce words [7]. The (column) word of a Young tableau is
obtained by listing its entries in the order from bottom to top in the left-most row, then from
bottom to top in the next-to-left-most row, continuing until the top entry of the right-most row
is listed. A plactic monomial is the result of substituting in the word the hopping operator ai
for the number i. For example, the plactic monomials of the Young tableaux in (8) are
a2a1a2 a2a1a1
a3a2a2 a2a1a3 a3a1a2 a3a1a1
a3a2a3 a3a1a3
On Affine Fusion and the Phase Model 5
Acting on the state |µ〉 with the plactic monomials of Young tableaux of a fixed shape λ is
equivalent to using the Littlewood–Richardson rule to calculate the decomposition of the tensor
product L(λ)⊗ L(µ).
Now, if a triple tensor product L(κ)⊗L(λ)⊗L(µ) is considered, the procedure is not unique.
Most straightforwardly, calculating L(λ) ⊗ L(µ) first leads to a set of mixed tableaux, one for
each irreducible highest-weight representation in the decomposition. If the mixed tableaux are
replaced by Young diagrams of the same shape, then the result can be calculated with the rule
already described, applied a second time: L(κ)⊗ (L(λ)⊗ L(µ)).
On the other hand, one can also multiply the Young tableaux of shape κ and those of
shape λ, to obtain a new set of Young tableaux. These product Young tableaux can then be
added in the usual way to the Young diagram of shape µ, to obtain the desired decomposition
(L(κ)⊗L(λ) )⊗L(µ). The required product • of Young tableaux is described by the “bumping”
process [7]. Fundamental examples are
j k • i = i k
j
, i < j ≤ k;
i k • j = i j
k
, i ≤ j < k. (10)
If the Young tableaux are translated into words, these bumping identities are translated into
relations for the corresponding plactic algebra. For example, with j = k = i + 1, (10) yields
a2
i+1ai = ai+1aiai+1; and with j = k − 1 = i, we get aiai+1ai = ai+1a
2
i . The full relations of (6)
guarantee that performing the Young tableaux calculations instead with the plactic monomials,
in a fully algebraic way, will yield equivalent results.
3 Phase model: Yang–Baxter equation and Bethe ansatz
We now write the Yang–Baxter equation for the phase model and apply the algebraic Bethe
ansatz in a standard way, following [16]. For an elementary introduction to the algebraic Bethe
ansatz, see [17], and for a comprehensive treatment, consult [13], for example.
First, introduce an auxiliary space isomorphic to C2, and work in C2 ⊗H. Write(
α β
γ δ
)
:=
(
1 0
0 0
)
⊗ α+
(
0 1
0 0
)
⊗ β +
(
0 0
1 0
)
⊗ γ +
(
0 0
0 1
)
⊗ δ,
for α, β, γ, δ operators acting on H (or endomorphisms of H).
Use the creation and annihilation operators of the phase model to define a Lax matrix, or
L-operator on C2 ⊗H
Li(u) :=
(
1 uϕ†i
ϕi u
)
, (11)
where u is the spectral parameter. The monodromy matrix is then
M(u) = Ln(u)Ln−1(u) · · ·L1(u) =
(
A(u) B(u)
C(u) D(u)
)
,
where the last equality just establishes the standard notation. For the simple L-operator of (11),
one finds B(u) = D(u)ϕ†n and C(u) = ϕnA(u). For n = 3 we find
A(u) = 1 + u(ϕ†2ϕ1 + ϕ†3ϕ2) + u2ϕ†3ϕ1, D(u) = u3 + u2(ϕ3ϕ
†
2 + ϕ2ϕ
†
1) + uϕ3ϕ
†
1. (12)
The monodromy matrix satisfies the fundamental relation
R12(u/v)M1(u)M2(v) = M2(v)M1(u)R12(u/v), (13)
6 M.A. Walton
with R-matrix given by
R(x) =
x
x−1 0 0 0
0 0 x
x−1 0
0 1
x−1 1 0
0 0 0 x
x−1
.
This relation works on H extended by two copies of the auxiliary C2, and the indices 1 and 2
indicate which of them the operators implicate.
The R-matrix satisfies the quantum Yang–Baxter equation
R12(u/v)R13(u)R23(v) = R23(v)R13(u)R12(u/v). (14)
On the other hand, (13) defines the so-called quantum Yang–Baxter algebra, satisfied by the
entries A(u), B(u), C(u), D(u) of the monodromy matrix. For example, we find
[B(u), B(v)] = 0, (15)
important here. The commuting B(u) will be used as creation operators for a basis of states in
the phase-model Hilbert space.
To demonstrate integrability, define the transfer matrix
T (u) := trM(u) = A(u) +D(u), (16)
where the trace is over the auxiliary space. Equation (13) guarantees that
[T (u), T (v)] = 0, (17)
so that T (u) =
∑
r u
rTr is the generating function of integrals of motion: [Tr, Ts] = 0. The
Hamiltonian of the phase model is recovered as
H = −1
2
(T1 + Tn−1) = −1
2
n∑
i=1
(
ϕiϕ
†
i+1 + ϕ†iϕi+1
)
.
By (12), the transfer matrix is
T (u) = 1 + u
(
ϕ†2ϕ1 + ϕ†3ϕ2 + ϕ†1ϕ3
)
+ u2
(
ϕ†3ϕ1 + ϕ†2ϕ3 + ϕ†1ϕ2
)
+ u3
= 1 + u(a2 + a3 + a1) + u2(a3a2 + a2a1 + a1a3) + u3.
The general result [16] is
T (u) =
n∑
r=0
urer(A), (18)
where er(A) indicates the r-th cyclic elementary symmetric polynomial, the sum of all cyclically
ordered products of r distinct hopping operators ai
er(A) =
∑
|I|=r
∏
i∈I
ai. (19)
In a monomial ai1ai2 · · · air , the relative order of 2 operators aij , aik only matters if ij and ik
differ by 1 modn, because of (4). Suppose the nodes of the circular affine su(n) Dynkin diagram
are numbered from 1 to n in clockwise fashion. Then anticlockwise cyclic ordering specifies
On Affine Fusion and the Phase Model 7
that if ik = ij+1 modn, then aij occurs to the right of aik . Thus, for n = 3, a3a2 is anticlockwise
circularly ordered, while a2a3 is not – the latter is connected to the “long way” anticlockwise
around the circular Dynkin diagram.
By (17), we know that
[er(A), er′(A)] = 0. (20)
For n = 3, the only non-trivial relation is
[e1(A), e2(A)] = [a1 + a2 + a3, a3a2 + a2a1 + a1a3] = 0.
Rewriting the defining relations (5) as
[a3, a3a2] = [a3, a1a3] = [a2, a2a1] = [a2, a3a2] = [a1, a1a3] = [a1, a2a1] = 0
makes obvious that it is satisfied.
The integrals of motion er(A) are related to Schur polynomials. By the substitution ai → xi,
one recovers the elementary symmetric polynomials er(x) = sΛr(x), the Schur polynomials for
the fundamental su(n) representations. The definition (19) therefore produces noncommutative
Schur polynomials for the fundamental representations of su(n). The integrability result (20)
allows us to define noncommutative Schur polynomials for all su(n) representations. Since
the er(A) commute, the Jacobi–Trudy formula
sλ(A) = det
(
eλti−i+j(A)
)
(21)
makes sense. Here λti is the i-th integer of λt, the transpose of the partition specifying λ.
For example, with n = 3 and λ = Λ1 + Λ2, we find
sΛ1+Λ2(A) = det
(
e2(A) e3(A)
e0(A) e1(A)
)
= det
(
a3a2 + a2a1 + a1a3 1
1 a1 + a2 + a3
)
= a2a
2
1 + a1a3a1 + a3a
2
2 + a2a1a2 + a3a2a3 + a1a
2
3
+ (a3a2a1 + a1a3a2 + a2a1a3 − 1). (22)
The terms of vanishing weight are enclosed in brackets.
Furthermore, the integrability (20) implies that the noncommutative Schur polynomials com-
mute among themselves4
[sλ(A), sµ(A)] = 0. (23)
One can therefore hope to find a basis diagonal in all these operators.
For that to be possible, the so-called Bethe ansatz equations must be satisfied [13, 17]. In
more detail, the Bethe state (or vector) uses the commuting operators B(u) as creation operators
to construct basis elements from the vacuum
|b(x)〉 := B
(
x−1
1
)
B
(
x−1
2
)
· · ·B
(
x−1
k
)
|0〉, (24)
that depend on invertible indeterminates x = (x1, . . . , xk). Since [N,B(u)] = B(u), each opera-
tor B(x−1
i ) injects a particle, and the states (24) have level k. Recalling (15), we see that |b(x)〉
4Terminology aside, it may be surprising that the noncommutative Schur polynomials commute. It was shown
in [6], however, that the case studied here is but one of a more general class of such noncommutative Schur
polynomials, that commute among themselves. The noncommutative arguments need only satisfy relations that
are implied by those in (4) but do not themselves imply (4).
8 M.A. Walton
is completely symmetric in the variables x−1
1 , . . . , x−1
k . This can be made completely explicit
using level-k symmetric polynomials [16]
|b(x)〉 =
∑
λ∈Pk
+
sλt
(
x−1
1 , . . . , x−1
k
)
|λ〉. (25)
Now the Bethe vector |b(x)〉 can be shown to be an eigenvector of the transfer matrix (16)
T (u)|b(x)〉 =
{[
1 + (−1)kek(x)un+k
] k∏
i=1
1
1− uxi
}
|b(x)〉, (26)
using the Yang–Baxter algebra (which follows from the fundamental relation (13)) and properties
of the 0-particle vacuum |0〉 [16]5. But this works only if x obeys the Bethe ansatz equations
xn+k
1 = · · · = xn+k
k = (−1)k−1x1x2 · · ·xk. (27)
Remarkably, the solutions to (27) are in 1-1 correspondence with weights in P k+. To see
roughly how this works, consider the variables yi = x−1
i xi+1, with indices defined cyclically
mod k. Think of a pie that can be divided into n + k equal portions of angles 2π/(k + n) [1].
Each yi is an (n+ k)-th root of unity, and so determines a slice with a number of portions, the
slice size. Since y1y2 · · · yk = 1, each x = (x1, . . . , xk) determines a slicing of the pie into k slices,
or a k-slicing. Furthermore, there is an n-slicing complementary to each k-slicing: where the pie
is cut and where it is not cut are interchanged. The slice sizes in the n-slicing give the Dynkin
labels of shifted weights σ + ρ and thus the weights σ ∈ P k+. The solutions to the Bethe ansatz
equations can therefore be labelled by these σ ∈ P k+: x = xσ.
For complete detail, see [16]. The result, valid for all n and k, is that the solutions to the
Bethe ansatz equations, or Bethe roots, are in 1-1 correspondence with weights in P k+.
4 Affine fusion
With the Bethe ansatz equations satisfied at x = xσ, so is equation (26). Then |b(xσ)〉 is an
eigenvector of the transfer matrix, and an eigenvector of all the er(A), in view of (18). The
eigenvalues can be determined from (26), and one finds
er(A)|b(xσ)〉 = hr(xσ)|b(xσ)〉,
where hr(x) is the r-th complete symmetric polynomial. The noncommutative Jacobi–Trudy
formula (21) then implies
sλ(A)|b(xσ)〉 = det
(
hλti−i+j(xσ)
)
|b(xσ)〉 = sλt(xσ)|b(xσ)〉. (28)
The last equality follows from a well-known identity for symmetric polynomials, an alternative,
dual Jacobi–Trudy formula.
The connection with affine fusion now becomes clear, because
sλt(xσ) =
Sλ,σ
SkΛn,σ
. (29)
Here Sλ,σ denotes an element of the unitary modular S-matrix [11] for su(n)k. Since
(k)Nν
λ,µ =
∑
κ∈Pk
+
Sλ,κSµ,κSν∗,κ
SkΛn,κ
(30)
5In (26), ek(x) = x1 · · ·xk is the k-th elementary symmetric polynomial.
On Affine Fusion and the Phase Model 9
by the Verlinde formula [18], the fusion eigenvalues Sλ,σ/SkΛn,σ obey(
Sλ,σ
SkΛn,σ
)(
Sµ,σ
SkΛn,σ
)
=
∑
ν∈Pk
+
(k)Nν
λ,µ
(
Sν,σ
SkΛn,σ
)
.
Therefore, (28) and (29) combine into
sλ(A)|b(xσ)〉 =
Sλ,σ
SkΛn,σ
|b(xσ)〉,
so that
sλ(A)sµ(A)|b(xσ)〉 =
∑
ν∈Pk
+
(k)Nν
λ,µ
(
Sν,σ
SkΛn,σ
)
|b(xσ)〉 =
∑
ν∈Pk
+
(k)Nν
λ,µsν(A)|b(xσ)〉.
The fusion algebra is commutative, (k)Nν
λ,µ = (k)Nν
µ,λ. It is significant that the commutativity
is guaranteed here by integrability: the noncommutative Schur polynomials commute by (23)
because they are integrals of motion, existing due to the Yang–Baxter equation (14).
By the Bethe ansatz equations, the Bethe vectors |b(xσ)〉 for σ ∈ P k+ form a complete or-
thogonal (but not normalized) basis of the Hilbert space at level k. Going back to (25), we can
relate them to the standard basis
|b(xσ)〉 =
∑
λ∈Pk
+
sλt
(
x−1
σ
)
|λ〉 =
∑
λ∈Pk
+
S∗λ,σ
SkΛn,σ
|λ〉.
The unitarity of the modular S-matrix then yields∑
σ∈Pk
+
SkΛn,σSσ,µ|b(xσ)〉 = |µ〉,
and then applying sλ(A) leads to
sλ(A)|µ〉 =
∑
ν∈Pk
+
(k)Nν
λ,µ|ν〉, (31)
taking the Verlinde formula into account.
Since (k)Nν
λ,kΛn = δνλ (the highest weight kΛn labels the identity field), we find
sλ(A)|kΛn〉 = |λ〉.
This is highly reminiscent of the state-field correspondence in conformal field theory (see [4]),
hinting that the operators sλ(A) play the role in the phase model of the primary fields in the
corresponding WZNW model.
This becomes clear, however, when
〈ν|sλ(A)|µ〉 = (k)Nν
λ,µ ∀ ν, µ ∈ P k+ (32)
and
(k)Nλ,µ,ν = 〈kΛn|sλ(A)sµ(A)sν(A)|kΛn〉
are written. Indeed, the noncommutative Schur polynomials play the role of primary fields, for
any number of them
(k)Nλ1,λ2,...,λN = 〈λ∗1|sλ2(A) · · · sλN−1
(A)|λN 〉 = 〈kΛn|sλ1(A) · · · sλN (A)|kΛn〉.
10 M.A. Walton
Like the noncommutative Schur polynomials, the Bethe vectors have an affine-fusion-algebraic
significance [15]. Define the (non-normalized) vectors
|bσ〉 =
|b(xσ)〉
〈b(xσ)|b(xσ)〉
= SkΛn,σ
∑
λ∈Pk
+
S∗λ,σ|λ〉.
Then by Verlinde’s formula (30), the formal fusion product ∗ yields
|λ〉 ∗ |µ〉 =
∑
ν∈Pk
+
(k)Nν
λ,µ|ν〉 ⇒ |bσ〉 ∗ |bτ 〉 = δσ,τ |bσ〉.
That is, rescaled Bethe vectors are the idempotents of the affine fusion algebra.
5 New perspective
Affine fusion appears in other integrable models – see [9], for early examples. The simple
realization afforded by the phase model [16], however, provides a fresh, new perspective on the
old subject. In this section we start to exploit it.
5.1 Threshold level
Perhaps the most striking property of the central result (31) is how the level-dependence of
fusion is realized. The noncommutative Schur polynomial sλ(A) has no dependence on the
level! At the price of noncommutativity, the same sλ(A) works for all levels k. In the expression
sλ(A) |µ〉, all level-dependence lies in the state |µ〉, a much simpler object.
Affine fusion has a simple dependence on the level, described well by the concept of a thresh-
old level [3, 12]. Each highest weight representation in the decomposition of a fusion will appear
at all levels greater than or equal to a minimum, non-negative integer value. This threshold
level is best understood as a consequence of the Gepner–Witten depth rule [8], or a refinement
thereof, conjectured in [12] and proved in [5].
All possible fusion decompositions can be given simply by treating the level as a variable, and
writing multi-sets of threshold levels as subscripts. For example, we rewrite the su(3) tensor
product decomposition (9) as
L
(
Λ1 + Λ2
)⊗2
↪→ L(0)2 ⊕ 2L
(
Λ1 + Λ2
)
2,3
⊕ L
(
3Λ1
)
3
⊕ L
(
3Λ2
)
3
⊕ L
(
2Λ1 + 2Λ2
)
4
. (33)
A multi-set of threshold levels can be replaced by a threshold polynomial T (t)νλ,µ with non-
negative integer coefficients [10]; so we can also write
L
(
Λ1+ Λ2
)⊗2
↪→ t2L(0)⊕
(
t2 + t3
)
L
(
Λ1+ Λ2
)
⊕ t3L
(
3Λ1
)
⊕ t3L
(
3Λ2
)
⊕ t4L
(
2Λ1+ 2Λ2
)
.
In general, the threshold polynomials are
T (t)νλ,µ =
∞∑
t′
(t′)nνλ,µt
t′ .
Here the threshold multiplicities (t)nνλ,µ satisfy
(k)Nν
λ,µ =
k∑
t
(t)nνλ,µ,
On Affine Fusion and the Phase Model 11
so that T (1)νλ,µ = (∞)Nν
λ,µ = T νλ,µ, the tensor-product multiplicities. We also find
(k)nνλ,µ = (k)Nν
λ,µ − (k−1)Nν
λ,µ, (34)
where we have put (k−1)Nν
λ,µ = 0 if any of λ, µ, ν are not in P k−1
+ .
In a similar way, the level-dependence can be incorporated into (31) simply by using |µ〉 with
variable level. The fusion decomposition (33) can be derived easily this way by applying (22) to
|Λ1 + Λ2 + (k − 2)Λ3〉, for example.
More generally, write µ̄ = µ1Λ1 + · · ·+ µn−1Λn−1 and define µ̄k := µ̄+ (k − µ1 − µ2 − · · · −
µn−1)Λn. Then
sλ(A)|µ̄k〉 =
∑
ν̄∈Pk
+
∑
t≤k
(t)nνλ,µ|ν̄k〉. (35)
In the limit of large k, the tensor product is recovered, and (35) becomes
sλ(A)|µ̄∞〉 =
∑
ν̄∈Pk
+
∞∑
t
(t)nνλ,µ|ν̄∞〉. (36)
Since sλ(A) does not depend on the level of λ, so that λ → λ̄∞ doesn’t change anything, this
justifies our notation
L(λ̄)⊗ L(µ̄) ↪→
⊕
ν̄∈P+
T ν̄λ̄,µ̄(t)L(ν̄). (37)
Another advantage of the phase-model realization of affine fusion is that, unlike in the WZNW
model, the level is not fixed – it is just the total particle number. Changes in level can be
described in a simple, algebraic way by the operators ϕ†i , ϕi of the phase algebra (1). In [16],
recursion relations involving fusion multiplicities at levels k and k + 1 were derived using this
observation. Such relations are difficult to see in other ways6.
Let us treat the threshold multiplicities (34) in similar spirit. Notice that ϕ†n|µ̄k−1〉 = |µ̄k〉.
So we calculate
[sλ(A), ϕ†n]|µ̄k−1〉 =
∑
ν∈Pk
+
(k)Nν
λ,µ|ν̄k〉 −
∑
ν∈Pk−1
+
(k−1)Nν
λ,µϕn|ν̄k−1〉.
So the phase-model version of (34) is
〈ν̄k|[sλ(A), ϕ†n]|µ̄k−1〉 = (k)nνλ,µ. (38)
Once a particular noncommutative Schur polynomial sλ(A) is calculated, the interesting
operator [sλ(A), ϕ†n] is easy to write down, since [ai, ϕ
†
n] = δi,1ϕ
†
1πn.
5.2 Higher-genus Verlinde dimensions
As another new application of the phase-model realization of su(n)k affine fusion, we consider
higher-genus fusion, i.e. higher-genus Verlinde dimensions [18].
In the WZNW model, the fusion multiplicity (k)Nν
λ,µ is also the dimension of the space of
conformal blocks for the corresponding 3-point function, its Verlinde dimension. The conformal
blocks originate from correlation functions on the sphere with 3 points marked by the 3 primary
6See what were called “identities of the Feingold type” in [19], however, which relate fusion multiplicities at
different levels.
12 M.A. Walton
fields, and the fusion multiplicity can be represented graphically by a 3-legged vertex that arises
in a degenerate limit of the marked sphere.
A sphere with n marked points corresponds to a trivalent fusion graph with no loops. But
higher-genus Riemann surfaces can also be considered, and so fusion graphs with loops are
allowed. For such higher-genus Riemann surfaces with marked points, the trivalent graph that
results is not unique. The conformal bootstrap, however, ensures that the Verlinde dimension
calculated from any of the graphs is the same. So, all dimensions can be built from the genus-0
3-point ones, for example. By this reasoning, one can see that the Verlinde formula extends
to [18]
(k,g)Nλ1,λ2,...,λN =
∑
σ∈Pk
+
(SkΛn,σ)2(1−g)
(
Sλ1,σ
SkΛn,σ
)
· · ·
(
SλN ,σ
SkΛn,σ
)
.
Here the left-hand side indicates the su(n)k Verlinde dimension for a genus-g Riemann sphere
with N marked points.
In the phase-model realization, the argument above again applies, so that we can build all
the required Verlinde dimensions from (32). So, for example,
(k,1)Nλ1,λ2 =
∑
α,β∈Pk
+
(k,0)N
λ∗1
α∗,β
(k,0)Nβ
α,λ2
=
∑
α,β∈Pk
+
〈λ∗1|sα∗ |β〉〈β|sα|λ2〉 = 〈λ∗1|
∑
α∈Pk
+
sα∗sα|λ2〉.
Here we have dropped the arguments from the noncommutative Schur polynomials, α∗ indicates
the weight charge-conjugate to α, e.g., and we have used the completeness of the standard basis
states.
Using this genus-1, 2-point function, the general Verlinde dimension can be constructed, with
the nice result
(k,g)Nλ1,...,λN = 〈λ∗1|
∑
α∈Pk
+
sα∗sα
g
sλ2 · · · sλN−1
|λN 〉
= 〈kΛn|
∑
α∈Pk
+
sα∗sα
g
sλ1sλ2 · · · sλN |kΛn〉. (39)
Recall that all the noncommutative Schur polynomials commute. Notice that
∑
α∈Pk
+
sα∗sα can be
interpreted as a genus-generating operator, or handle-creation operator.
6 Conclusion
Let us first point out the new results obtained. The existence of a threshold level for su(n)k affine
fusion is made plain in the phase-model realization. The noncommutative Schur polynomials
do not depend on the level; all dependence on k lies in the basis vectors |λ〉. The threshold-
polynomial notation (37) was validated easily in the phase-model realization by (36). It was
also shown in (38) how threshold multiplicities may be calculated using, in addition to the
noncommutative Schur polynomials of the hopping (affine local plactic) algebra, the creation
operator ϕ†n of the phase algebra.
The remarkable result (32) of [16] was generalized to the elegant formula (39) for arbitrary
Verlinde dimensions, at any genus g and for any number N of marked points.
On Affine Fusion and the Phase Model 13
Most of this paper is not original, however. The bulk of it was devoted to a non-rigorous
summary of the integrable, phase-model realization of affine su(n) fusion discovered recently
by Korff and Stroppel [16] (also reviewed in [14, 15]). The goal was to provide a brief, easily
accessible treatment in the hope of interesting others in this nice work. I believe that the Korff–
Stroppel integrable realization of affine fusion will help us understand better affine fusion, the
WZNW models and perhaps more general rational conformal field theories.
Acknowledgements
I thank Elaine Beltaos, Terry Gannon, Ali Nassar and Andrew Urichuk for discussions and/or
reading the manuscript. This research was supported in part by a Discovery Grant from the
Natural Sciences and Engineering Research Council of Canada.
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http://dx.doi.org/10.1139/p94-067
1 Introduction
2 Phase model: Hilbert space and operator algebras
3 Phase model: Yang-Baxter equation and Bethe ansatz
4 Affine fusion
5 New perspective
5.1 Threshold level
5.2 Higher-genus Verlinde dimensions
6 Conclusion
References
|
| id | nasplib_isofts_kiev_ua-123456789-148697 |
| institution | Digital Library of Periodicals of National Academy of Sciences of Ukraine |
| issn | 1815-0659 |
| language | English |
| last_indexed | 2025-12-07T15:26:03Z |
| publishDate | 2012 |
| publisher | Інститут математики НАН України |
| record_format | dspace |
| spelling | Walton, M.A. 2019-02-18T17:56:45Z 2019-02-18T17:56:45Z 2012 On Affine Fusion and the Phase Model / M.A. Walton // Symmetry, Integrability and Geometry: Methods and Applications. — 2012. — Т. 8. — Бібліогр.: 19 назв. — англ. 1815-0659 2010 Mathematics Subject Classification: 81T40; 81R10; 81R12; 17B37; 17B81; 05E05 DOI: http://dx.doi.org/10.3842/SIGMA.2012.086 https://nasplib.isofts.kiev.ua/handle/123456789/148697 A brief review is given of the integrable realization of affine fusion discovered recently by Korff and Stroppel. They showed that the affine fusion of the su(n) Wess-Zumino-Novikov-Witten (WZNW) conformal field theories appears in a simple integrable system known as the phase model. The Yang-Baxter equation leads to the construction of commuting operators as Schur polynomials, with noncommuting hopping operators as arguments. The algebraic Bethe ansatz diagonalizes them, revealing a connection to the modular S matrix and fusion of the su(n) WZNW model. The noncommutative Schur polynomials play roles similar to those of the primary field operators in the corresponding WZNW model. In particular, their 3-point functions are the su(n) fusion multiplicities. We show here how the new phase model realization of affine fusion makes obvious the existence of threshold levels, and how it accommodates higher-genus fusion. This paper is a contribution to the Special Issue “Superintegrability, Exact Solvability, and Special Functions”. The full collection is available at http://www.emis.de/journals/SIGMA/SESSF2012.html. I thank Elaine Beltaos, Terry Gannon, Ali Nassar and Andrew Urichuk for discussions and/or reading the manuscript. This research was supported in part by a Discovery Grant from the Natural Sciences and Engineering Research Council of Canada. en Інститут математики НАН України Symmetry, Integrability and Geometry: Methods and Applications On Affine Fusion and the Phase Model Article published earlier |
| spellingShingle | On Affine Fusion and the Phase Model Walton, M.A. |
| title | On Affine Fusion and the Phase Model |
| title_full | On Affine Fusion and the Phase Model |
| title_fullStr | On Affine Fusion and the Phase Model |
| title_full_unstemmed | On Affine Fusion and the Phase Model |
| title_short | On Affine Fusion and the Phase Model |
| title_sort | on affine fusion and the phase model |
| url | https://nasplib.isofts.kiev.ua/handle/123456789/148697 |
| work_keys_str_mv | AT waltonma onaffinefusionandthephasemodel |