Prepared for submission to JHEP

Surprises from the resummation of ladders in the

ABJ(M) cusp anomalous dimension

Marisa Boninia, Luca Griguoloa, Michelangelo Pretia and Domenico Seminarab

a Dipartimento di Fisica e Scienze della Terra, Universita di Parma and INFN Gruppo Collegato

di Parma, Viale G.P. Usberti 7/A, 43100 Parma, Italyb Dipartimento di Fisica, Universita di Firenze and INFN Sezione di Firenze, via G. Sansone 1,

50019 Sesto Fiorentino, Italy

E-mail: marisa.bonini@fis.unipr.it, luca.griguolo@fis.unipr.it,

michelangelo.preti@fis.unipr.it, domenico.seminara@fi.infn.it

Abstract: We study the cusp anomalous dimension in N = 6 ABJ(M) theory, identifying

a scaling limit in which the ladder diagrams dominate. The resummation is encoded into

a Bethe-Salpeter equation that is mapped to a Schroedinger problem, exactly solvable due

to the surprising supersymmetry of the effective Hamiltonian. In the ABJ case the solution

implies the diagonalization of the U(N) and U(M) building blocks, suggesting the existence

of two independent cusp anomalous dimensions and an unexpected exponentiation struc-

ture for the related Wilson loops. While consistent with previous perturbative analysis,

the strong coupling limit of our result does not agree with the string theory computation,

emphasizing a difference with the analogous resummation in the N = 4 case.

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Contents

1 Introduction and summary of the results 1

2 The 1/2 BPS generalized cusped Wilson line in ABJ(M) theory 5

2.1 The 1/2 BPS Wilson line 5

2.2 The generalized cusp 6

3 The cusp anomalous dimension in ABJ(M) theory and its computation

through ladder diagrams 7

3.1 The cusp anomalous dimension in ABJ(M) theory 7

3.2 The scaling limit selecting ladder diagrams 9

4 Bethe-Salpeter equation for the generalized cusp in ABJ(M) at leading

order in the scaling limit 10

4.1 General solution in d = 3 13

4.2 The straight-line limit: the solution for ε 6= 0 15

5 The determination of Γcusp(ϕ) 17

6 Conclusions 19

A N = 6 three-dimensional Chern-Simons-matter theory 21

B Useful integral 22

1 Introduction and summary of the results

The duality between N = 6 Super Chern-Simons theory with matter (also known as

ABJ(M) theory [1, 2]) and string theory on AdS4 × CP3 represents one the most inter-

esting possibilities to explore AdS/CFT correspondence beyond the original paradigm.

Although it seems to share many similarities with the cousin N = 4 Super Yang-Mills

theory in four dimensions, there are still many aspects calling for a better comprehension.

– 1 –

Supersymmetric Wilson loops, in particular, provide a rich class of BPS observables [3]-[9]

that can be computed exactly through localization technique in the simplest situations [10].

While their quantum behavior is still rather mysterious in the general case [11], the well-

understood 1/2 BPS and 1/6 BPS circles exhibit an intriguing non-trivial interpolation

between weak and strong coupling regime [12]-[15]. A careful study of the relevant matrix

models [15] has also unveiled a variety of phenomena of string/M-theory origin [16]-[18].

This contrasts with the relative simplicity of N = 4 SYM, where a gaussian matrix model

[19, 20] describes exactly the dynamics of 1/2 BPS Wilson loops [21].

Wilson loops are also relevant in gauge theories (supersymmetric or not) because they

encode important properties of scattering amplitudes and infrared radiation: the cusp

anomalous dimension Γ(ϕ), originally introduced in [22] as the ultraviolet divergence of a

Wilson loop with Euclidean cusp angle ϕ, appears in fact in many interesting physical situ-

ations. It was computed in QCD to two-loop order in [23] and, in light-like limit ϕ→ i∞,

provides crucial universal informations [24]. In supersymmetric theories Γ(ϕ) is not a BPS

observable, making its exact computation in N = 4 SYM a difficult challenge. In the

light-like limit [25] integrability controls its all-order behavior through an integral equa-

tion, whose solution matches both weak coupling expansions [26] and string computations

at the strong coupling [27]. In the general case a strategy [28, 29] for computing Γ(ϕ) was

later proposed (see [30–32] for the original introduction of the related TBA approach in

the case of local operators). The cusp anomalous dimension can be generalized including

an R-symmetry angle θ that distinguishes the coupling of the scalars to the two halves of

the cusp [38]. The new observable Γ(ϕ, θ) interpolates between BPS configurations and

generalized quark-antiquark potentials. Exact equations can be written applying integra-

bility and they have been checked successfully at three loops [28, 39]. In the near-BPS limit

it is possible to use localization to obtain the exact form of the so-called Bremsstrahlung

function [40, 41], that has been later directly recovered from the TBA equations [42, 43].

More recently localization has been used to derive the Bremsstrahlung function in presence

of local operator insertions [44] and in more general superconformal field theories [45].

As anticipated before, even for the simple circular 1/2 BPS Wilson loop the weak/strong

interpolation for ABJ(M) theory is not non-trivial. Integrability itself has been explored

here in a somehow limited range of situations [46]-[50]: when established it still depends

on an elusive interpolating function, h(λ) [51]-[53]. Recently a proposal for the functional

form of h(λ) has been advanced [54] and checked at two-loop level in string theory [55],

under suitable assumptions. Alternatively h(λ) could be determined by computing exactly

some quantity by integrability and confronting with the same calculation by localization

(or by other QFT techniques in which unknown functions are absent). A natural candidate

would be the ABJM Bremsstrahlung function, for which all order proposals exist [56], but

integrability has not been yet applied to its determination.

For this reason we think it is important to study Γ(ϕ, θ) in N = 6 Super Chern-Simons

theory: it would be useful to obtain exact QFT results to be compared with the integrability

approach and, at strong coupling, with string theory. Γ(ϕ, θ) has been introduced in

– 2 –

ABJ(M) theory in [57], where its computation at two-loop has been presented and its

exponentiation properties discussed. The two halves of the cusp are locally 1/2 BPS and

therefore couple also directly to the fermions of the theory, as originally discovered in [7],

and not only to the gauge connections and scalars. The resulting cusped Wilson loop is

not globally supersymmetric and its exact evaluation seems very challenging. Fortunately,

the analogous system in N = 4 SYM can be tackled in a particular limit through Feynman

diagrams resummation. In [39] a new scaling limit involving the complexified angle θ was

introduced,

iθ � 1, λ� 1, λ = λ exp(iθ/4) fixed. (1.1)

Here λ = g2N is the ’t Hooft coupling of N = 4 SYM. In this limit the leading order

contribution is simply given by ladder diagrams, where the rungs are made by scalar

exchanges. The ladder diagrams can be summed up efficiently using a Bethe-Salpeter

equation, solved exactly in the small ϕ limit. The strong coupling behavior has been

also examined, finding agreement with the corresponding string theory calculation [39].

Later it was also performed an analysis at next-to-leading order, generalizing the original

Bethe-Salpeter equation and computing the relevant corrections at strong coupling [58].

Remarkably these corrections have been also obtained from string theory and successfully

compared each other. As repeatedly stressed in the original analysis [39], the matching of

the strong coupling limit of the Bethe-Salpeter solution with the string theory computation

is quite surprising. The ladders limit, λ → 0 with λ fixed, is different from the strong

coupling limit λ→∞ with iθ � 1 fixed and the result could, in principle, depend on their

order: nevertheless they agree at leading and subleading level.

In this paper we consider a similar limit in three-dimensional ABJ(M) theory,

iθ � 1, λr � 1, λr = λr cosθ

2, (r = 1, 2) (1.2)

obtaining some exact results for Γ(ϕ, λi). Here λ1 = Nk , λ2 = M

k are the ’t Hooft couplings

of the ABJ(M) theory with gauge group U(N)⊗U(M), while k is the Chern-Simons level.

The presence of fermionic couplings to the cusped loop inherits a surprising supersymmetric

structure in the relevant Bethe-Salpeter equation. More precisely, the effective Schroedinger

problem, associated to the integral equation that resums planar diagrams in N = 6 Super

Chern-Simons, enjoys an unexpected quantum mechanical supersymmetry. Because only

the ground state matters in determining Γ(ϕ, λi) [39], supersymmetry produces an exact

expression for any value of the opening angle ϕ. This is in sharp contrast with the N = 4

case, where an analytic solution for the Bethe-Salpeter equation exists only at ϕ = 0 (that

in this case is the only supersymmetric point of the associated Schroedinger equation [39]).

In the ABJM case (N = M) we get a very simple solution: Γ(ϕ, λ) is exact at one-loop level,

as in an abelian theory. The delicate balance between bosonic and fermionic contributions,

encoded into the effective supersymmetric quantum mechanics, exponentiates without non-

abelian correction the one-loop term. As a matter of fact we do not observe any transition

between a weak-coupling and a strong-coupling regime and we cannot match our result

– 3 –

with semiclassical computations in string theory, suggesting that it should exist a problem

with the order of limits in this case.

In the ABJ case (N 6= M) the story is even more intriguing: the Bethe-Salpeter equa-

tion reduces to two coupled integral equations, resumming contributions from the upper

(N ×N) and the lower (M ×M) blocks of the holonomy of the U(N |M) superconnection

defining the Wilson loop [7]. By diagonalizing the system we end up with a two-dimensional

supersymmetric Schroedinger equation. From the knowledge of its ground state energy and

wave-function we reconstruct the original cusped Wilson loops. This works in full general-

ity in three dimensions with explicit UV an IR cut-offs on the lines, while using dimensional

regularization we get an explicit solution only for ϕ = 0. The main result of our investiga-

tion is that, in the ABJ case, the original cusped Wilson loop, the trace of the holonomy

of the superconnection defined in [7], does not renormalize multiplicatively but it mixes,

at quantum level, with the supertrace. As a consequence we end up with two indepen-

dent cusp anomalous dimensions, related to the renormalization constants of the operator

eigenstates (with respect to the mixing). On the other hand we could have expected this

fact from the very beginning: taking the trace of the holonomy of the superconnection is

required to preserve the global supersymmetry of the 1/2 BPS Wilson loop [7], In our case

supersymmetry is generically broken and we observe a mixing between the trace and (say)

the supertrace. Again the cusp anomalous dimensions are exact at one-loop level

Γ(1)cusp(ϕ) =

√λ1λ2

cos ϕ2,

Γ(2)cusp(ϕ) =−

√λ1λ2

cos ϕ2.

(1.3)

We expect that subleading corrections should change the simple pattern Γ(2)cusp = −Γ

(1)cusp:

we will come back on this point in the conclusions.

The paper is structured in the following way: in Section 2 we introduce the cusped Wil-

son loop in ABJ(M) theory as the trace of a superconnection, following [57]. The limit

in which the ladder diagrams dominate is described in Section 3, where we also discuss

how to derive the cusp anomalous dimensions both in the explicit cut-off scheme and in

dimensional regularization. In Section 4 we derive the relevant Bethe-Salpeter equation

and, after diagonalization, we obtain the associated Schroedinger equation. We solve the

supersymmetric Schroedinger equations for generic opening angle ϕ, in the cut-off scheme,

and for ϕ = 0 in dimensional regularization. In Section 5 we obtain the cusp anomalous

dimensions at leading order and discuss the operator mixing for the cusped Wilson loops.

Our conclusions and the future directions to improve our results appear in Section 6. Two

appendices complete our presentation.

– 4 –

2 The 1/2 BPS generalized cusped Wilson line in ABJ(M) theory

In this Section we review the construction of supersymmetric Wilson lines in ABJ(M)

theory [7]. In d = 3 the generalized gauge connection can be defined in two different ways

according to the degree of preserved supersymmetry. Indeed, we can consider a purely

bosonic gauge connection whose holonomy is, for a suitable choice of the path, 1/6 BPS.

On the other hand, adding on the lines local couplings to the fermions, we can interpret

the Wilson operator as the holonomy of a U(N |M) superconnection obtaining, for the

infinite straight line, a 1/2 BPS operator. Its peculiar structure was also investigated via

the so-called Higgsing procedure which gives a physical explanation for the appearance of

the superconnection [8].

2.1 The 1/2 BPS Wilson line

In ABJ(M) theory the gauge sector consists of two gauge fields (Aµ) ji and (Aµ) j

ibelonging

respectively to the adjoint of U(N) and U(M). The matter sector contains the complex

fields (CI)ii and (CI) i

ias well as the fermions (ψI)

ii

and (ψI) ii . The fields (C, ψ) transform

in the (N, M) of the gauge group U(N)×U(M) while the fields (C, ψ) live in the (N,M).

The additional capital index I = 1, 2, 3, 4 belongs to the R-symmetry group SU(4).

The central idea of [7] is to replace the U(N) × U(M) gauge connection with the super-

connection

L(τ) ≡ −i

iA

√2πκ |x|ηI ψI√

2πκ |x|ψI ηI iA

with

A ≡ Aµxµ − 2πiκ |x|M I

J CICJ

A ≡ Aµxµ − 2πiκ |x|M I

J CJCI ,

(2.1)

belonging to the super-algebra of U(N |M). In (2.1) the coordinates xµ(τ) define the

contour of the loop operator, while M IJ , M I

J are the scalar couplings and ηαI , ηIα are

fermionic (2-components Grassmann even quantities) ones.

The form of L is determined mainly by dimensional analysis and symmetry properties of

the fields. In d = 3 the scalars have classical dimension 1/2, so they could only appear as

bilinears, which are in the adjoint and therefore enter in the diagonal blocks together with

the gauge fields. Instead the fermions have dimension 1 and should appear linearly. Since

they transform in the bifundamental, they are naturally placed in the off-diagonal entries

of the matrix.

For a given a path C, it is possible to compute the holonomy of the superconnection (2.1)

W [C] ≡ P exp

(−i∫

CL(τ)dτ

). (2.2)

When the contour is a straight-line S, all the couplings can be chosen to be independent of τ ,

i.e. constant, in order to preserve the invariance under translations along the line. Further

restrictions on scalar and fermionic couplings follow from R−symmetry and supersymmetry

– 5 –

requirements. Actually, imposing δsusyL(τ) = 0 gives rise to loop operators which are

merely bosonic (η = η = 0) and at most 1/6 BPS [11]. The weaker condition of invariance

under supersymmetry up to a super-gauge transformation brings out the 1/2 BPS solution.

For (finite) closed path one has to carefully consider the boundary conditions obeyed by the

gauge functions to obtain a (super-)gauge invariant object. For instance, in the circle case

one has to take the trace of (2.2). For an infinite open circuit, such as the straight line, the

naive statement that the fields vanish when τ = ±∞ allows two possible supersymmetric

operators

W− =1

N −M Str

[Pexp

(−i∫dτL(τ)

)],

W+ =1

N +MTr

[Pexp

(−i∫dτL(τ)

)].

(2.3)

Usually one mainly considers the second possibility, since, for particular angles, it is con-

nected through a conformal transformation to a BPS closed loop [57]. Nevertheless the

supertraced holonomy has a crucial role as well.

2.2 The generalized cusp

x1

x2

Figure 1: The planar cusp

We consider the theory on the Euclidean space-time and take the contour depicted in

Figure 1. The two rays are in the plane (1, 2), intersect at the origin and are given by

xµ = {0, τ cosϕ

2, |τ | sin ϕ

2} −∞ ≥ τ ≤ ∞ . (2.4)

The angle between the rays is π − ϕ, thus ϕ = 0 gives the continuous straight line.

It is natural to consider scalar and fermionic couplings different on the two segments of the

cusp (but constant on each segment). The fermionic couplings have the factorized form

ηαiM = niMηαi and ηMiα = nMi ηiα , (2.5)

– 6 –

where the index i = 1, 2 specifies which edge of the cusp we are considering. As discussed

in [57] we can take:

η1α =(e−i

ϕ4 ei

ϕ4

), η1α = i

(eiϕ4

e−iϕ4

), (2.6)

and

η2α =(eiϕ4 e−i

ϕ4

), η2α = i

(e−i

ϕ4

eiϕ4

). (2.7)

The R−symmetry part of the couplings is totally unconstrained and we choose

n1M =(cos θ4 sin θ

4 0 0)

and n2M =(cos θ4 − sin θ

4 0 0), (2.8)

(and we denote by nMi the transpose of niM ). The matrices which couple the scalars on

the two edges are

M I1J = M I

1J =

− cos θ2 − sin θ2 0 0

− sin θ2 cos θ2 0 0

0 0 1 0

0 0 0 1

and M I

2J = M I2J =

− cos θ2 sin θ2 0 0

sin θ2 cos θ2 0 0

0 0 1 0

0 0 0 1

. (2.9)

The quantum holonomy of the super-connection L in a representation R of the supergroup

U(N |M) is by definition

〈WR〉 =1

dimR

∫D[A, A, C, C, ψ, ψ] e−SABJ TrR

[P exp

(−i∫

Γdτ L(τ)

)], (2.10)

where SABJ is the Euclidean action for ABJ(M) theory (see Appendix A). In the following

R will be chosen to be the fundamental representation of U(N |M): this implies that the

trace is taken in the fundamental representation N or M of the two gauge groups.

In general this Wilson loop operator is not supersymmetric unless θ = ±ϕ: in this case,

having chosen the trace in its definition, it can be mapped by a suitable conformal trans-

formation to a closed 1/6 BPS Wilson loop [57].

3 The cusp anomalous dimension in ABJ(M) theory and its computation

through ladder diagrams

In this Section we discuss the definition of the cusp anomalous dimension in ABJ(M)

theories and its computation in a limit in which ladder diagrams dominate. New features

in the N 6= M will emerge due to the exponentiation properties of the cusped loops.

3.1 The cusp anomalous dimension in ABJ(M) theory

We start by recalling the N = 4 SYM case: the generalized cusp anomalous dimension is

defined by the logarithmic divergent behaviour of a cusped Wilson loop [22]

〈Wcusp〉 ' e−Γcusp(ϕ,θ) logΛUVΛIR . (3.1)

– 7 –

Here ΛUV and ΛIR stand for the ultraviolet and infrared cut-offs respectively, that reg-

ularize the specific divergences associated to the cusp angle and the infinite extension of

the lines. Typically one takes ΛIR = 1/L, where L measures the (finite) length of the two

edges and ΛUV = 1/δ, with δ being a short-length scale either cutting or smoothing out

the the cusp singularity. Alternatively one can use dimensional regularization [23] and the

logarithm is replaced by a simple pole 1/ε. After the usual renormalization of the gauge

theory1, the relation between the bare and renormalized Wilson loop operator, for closed

contours, is

WBcusp = ZcuspWR

cusp , (3.2)

where the renormalization constant Zcusp depends on the dimensional regularization pa-

rameter ε and the subtraction point µ (in addition to the coupling constant λ). When

considering the standard cusp built out of two straight lines of length L, the relation (3.2)

must be, in general, corrected as follows [33–36]

WBcusp = ZcuspZopenWR

cusp. (3.3)

The second renormalization constant cancels spurious divergences due to the fact that we

are dealing with an open (and then non gauge-invariant) loop operator. The separation

between the two contributions is fixed by the renormalization condition Zcusp|ϕ=θ=0 = 1.

The cusp anomalous dimension is defined as2

Γcusp(ϕ, θ) = µd

dµlogZcusp , (3.4)

and plays the role of the anomalous dimension of a (non-local) quantum operator. Although

the open cusped Wilson loop is not gauge invariant, the cusp anomalous dimension Γcusp

turns out to be so. At the perturbative level the cusp divergence comes from diagrams

with propagators connecting both rays of the cusp and its exponentiation is governed

by their maximal non-abelian part [23]. In the N = 4 SYM case, the perfect balance

between the gauge and the scalars contributions cancels, in the Feynman gauge, all the

infinities related to integrations along the smooth part of the contour (Zopen = 1). Thus

only the singularities associated to the discontinuous behavior at the cusp appear and one

immediately singles out the relevant diagrams to be computed [37, 38].

In the ABJ(M) case the situation is more subtle. First of all the presence of the fermionic

contributions breaks that balance and divergences persist even in the straight-line limit,

at least in dimensional regularization and in the Landau gauge [57]. Moreover there are

two gauge groups and the fermionic interactions, that live in the off-diagonal sector of the

super-connection, mix non-trivially the U(N) and U(M) structures. Thus a non-standard

form of exponentiation for the divergence of the cusped Wilson loop is expected. Actually

in the ABJM case (N = M) the renormalization properties should be similar to the N = 4

SYM case, as explicitly checked at two-loop in [57]. In particular we will expect a single

exponential behavior for the vacuum expectation value of the cusped loop operator.

1In N = 4 SYM this step is superfluous, being the β-function vanishing2In the definition of the anomalous dimension, the limit ε → 0 is understood.

– 8 –

In the ABJ case (N 6= M), as usual in the theory of renormalization of composed local

operator, we would expect instead the arising of a matrix-valued set of renormalization

constants:

WBa = ZabWR

b , (3.5)

where a, b = ± refers to the traced and supertraced operators3. Similarly to (3.4), the

anomalous dimensions matrix is

(Γcusp)ab =

[µ∂

∂µlog Zcusp

]

ab

. (3.6)

In general the matrix Zcusp in (3.6) is obtained by rewriting the matrix Zab in (3.5) as

(ZopenZcusp)ab with normalization condition (Zcusp)ab|ϕ=θ=0 = δab. The scaling limit (1.2)

considered in this paper selects only diagrams connecting the two halves of the loop, then

Zopen = 1 and Z = Zcusp.

3.2 The scaling limit selecting ladder diagrams

In [39] it was considered the scaling limit (1.1) in N = 4 SYM: there pure scalar exchanges

between the rungs of the cusped Wilson loop become dominant and can be resummed by

means of a Bethe-Salpeter equation [39]. Subleading corrections can be also systemati-

cally included in this scheme and consistency at strong coupling with semiclassical string

computations has been found [58].

In our case, we want to consider a similar limit: upon a quick inspection of the perturbative

diagrams we recognize two types of relevant contributions (we refer to [57] for details on

the perturbative expansion and related computations). At one-loop we have the single

fermionic exchange (that is the same for the up and down diagonal blocks)

(c)(b)(a) (d) (e)

=

(2π

κ

)MN

M +N

Γ(

12 − ε

)

4π3/2−ε (µL)2ε 1

ε

cos θ2cos ϕ2

.

We have considered here the appropriate normalization of the trace. We notice that for

N = M this contribution is proportional to λ cos θ2 , suggesting to perform the scaling limit

iθ � 1, λ� 1, λ = λ cosθ

2fixed. (3.7)

The natural generalization to the case N 6= M is therefore

iθ � 1, λ1,2 � 1, λ1,2 = λ1,2 cosθ

2fixed. (3.8)

3These two combinations provides a natural gauge invariant basis for the information contained in the

superholonomy defined by L(τ). The presence of these two different possibilities is the source of all the

main differences with the N = 4 case, where only the trace makes sense.

– 9 –

At two-loop we observe that the above limit suppresses all the diagrams in which interac-

tions are present. Obviously the double-fermionic exchange survives but also a pure scalar

exchange comes into the game

(c)(b)(a) (d) (e)

Figure 4. Two loops bosonic diagrams: (a) One-loop corrected gauge propagators; (b) Correlator of two

composite scalar operators; (c) Correlator gauge field composite scalar operator; (d) Chern-Simons vertex

diagrams; (e) Gluon double exchange diagrams.

and it would correspond to a gauge transformation -albeit a singular one. In dimensional

regularization it yields a (✓,') independent pole in ✏ plus finite terms. Thus its contribution

to the divergent part of the cusp becomes ine↵ective when we impose the renormalization

condition discussed in subsec. 4.1. Consider now the other contribution in (5.3): as firstly

noted in [31] it possesses an unforeseen four-dimensional structure. When the two endpoints

lie on the same edge it is proportional to the the tree-level e↵ective propagator in N = 4

since Tr(M1M2) = 4 and thus it vanishes. When they lie instead on opposite edges it given

the following result

B(2) = �MN2

✓2⇡

◆2�2�

12 � ✏

�

4⇡3�2✏

✓cos'� cos2

✓

2

◆Z L

0d⌧1

Z 0

�Ld⌧2

1

((x � y)2)1�2✏, (5.4)

where the integral governing the divergence is the same of the four dimensional case when

we replace 2✏ with ✏.

Next we examine the graphs 4.(c), 4.(d) and 4.(e). The last one is identically zero for the

same reasons of the one-loop single exchange 3.(a). The diagram 4.(c) for the case of planar

loop was discussed in [31] where it was found to vanish. The same fate is shared by 4.(e)

as pointed out in [26]. Thus the only contribution originating from the bosonic diagram is

provided by (5.4).

5.2 Fermionic diagrams

Figure 5. One-loop corrected fermions

propagators

The simplest fermionic diagram appearing at the

second order in perturbation theory consists of

the exchange of the one-loop corrected fermion

propagator depicted in fig 5.

The one-loop two point function for the spinor

fields is briefly discussed in app. C. Remarkably

it again displays the four dimensional behaviour

already encountered in the bosonic case. Its form

in the DRED scheme is

D( I)

j

i(x)( J) l

k (y)E1 `oop

0= �i⇣

✓2⇡

◆� li�jk(N � M)

�2�

12 � ✏

�

16⇡3�2✏

1

((x � y)2)1�2✏. (5.5)

– 20 –

= −(

2π

κ

)2

MNΓ2(

12 − ε

)

16π3−2ε(µL)4ε cos2 θ

2

1

ε

ϕ

sinϕ.

This last contribution has exactly the same form of the one-loop scalar exchange in N = 4

SYM, except that here it appears at two-loop and the scaling behavior is different.

It is not difficult to realize that, at leading order, the generic diagrams surviving the limit

consist of ladders made by fermionic and scalar exchanges, that should therefore summed

up to obtain the complete result. We remark that the contributions coming from diagrams

ending on a single line, and so leading to divergences not related to the cusp renormalization

constant, are automatically suppressed in our limit. In the next Section we will derive an

efficient way to sum up all the relevant ladder diagrams.

4 Bethe-Salpeter equation for the generalized cusp in ABJ(M) at leading

order in the scaling limit

F

. . .

. . .

ST

= δab + F

. . .

. . .

. . .

. . .P

ST

Figure 2: Bethe-Salpeter equation at leading order

The general strategy to sum up ladder diagrams for Wilson loops constructed by straight-

lines was introduced in [59] and used in [39] to derive the leading behavior of the N = 4

SYM cusp anomaly in the scaling limit. We adapt their technique to our case.

We denote the sum of the ladder diagrams with end-points within the intervals (0, S) and

(0, T ) of the cusp by Fab(S, T ), where (a, b) are group indices that can be (i, j) ∈ U(N) or

(i, j) ∈ U(M), respectively. Fab(S, T ) satisfies a Bethe-Salpeter equation

Fab(S, T ) = δa

b +

∫ S

0ds

∫ T

0dt Fc

d(s, t)Pacdb(s, t) , (4.1)

– 10 –

that is shown schematically in Figure 2. In the scaling limit (1.2) the scalar and the

fermionic couplings of the loop dominate and one has to consider only exchanges of these

fields between the two segments of the Wilson loop. The indices sequence follows the path-

ordering of the Wilson loop and the scalar and fermionic propagators fix the kernel indices

as follow:

Pacdb(s, t) ' δa b δc d × (a “scalar” function of s and t), (4.2)

in particular

Piklj(s, t) = Mδi

jδk lP(B)(s, t) ,

Piklj(s, t) = δi

jδk lP(F )(s, t) ,

Piklj(s, t) = δi

jδk lP(F )(s, t) ,

Piklj(s, t) = Nδi

jδk lP(B)(s, t) ,

(4.3)

where P (F )(s, t) is the fermionic effective propagator and P (B)(s, t) the scalar effective

propagator (double exchange) defined by [57]:

P (F )(s, t) = −(

2π

k

)Γ(1/2− ε)µ2ε

4π3/2−εcos θ/2

cosϕ/2(∂s + ∂t)

1

(s2 + t2 + 2st cosϕ)12−ε ,

P (B)(s, t) =

(2π

k

)2 Γ2(1/2− ε)µ4ε

4π3−2ε

cos2 θ/2

cos2 ϕ/2

cos2 ϕ/2

(s2 + t2 + 2st cosϕ)1−2ε.

(4.4)

According to U(N) or U(M) indices (4.1) splits into

Fij(S, T ) = δi

j +

∫ S

0ds

∫ T

0dt(MFk

l(s, t)δijδk lP

(B)(s, t) + Fkl(s, t)δi

jδk lP(F )(s, t),

),

Fij(S, T ) = δi

j +

∫ S

0ds

∫ T

0dt(Fk

l(s, t)δijδk lP

(F )(s, t) +NFkl(s, t)δi

jδk lP(B)(s, t)

).

(4.5)

Thus, defining

F (S, T ) =1√N

TrN[Fij(S, T )] ,

F (S, T ) =1√M

TrM[Fij(S, T )] ,

(4.6)

we get

F (S, T ) =√N +

∫ S

0ds

∫ T

0dt(MNF (s, t)P (B)(s, t) +

√MNF (s, t)P (F )(s, t)

),

F (S, T ) =√M +

∫ S

0ds

∫ T

0dt(√

MNF (s, t)P (F )(s, t) +MNF (s, t)P (B)(s, t)).

(4.7)

– 11 –

F

. . .

. . .

=√N +MN · F

. . .

. . .

+√MN · F

. . .

. . .

F

. . .

. . .

=√M+√MN · F

. . .

. . .

+ MN · F

. . .

. . .

Figure 3: Bethe-Salpeter equation at leading order

Changing variables according to s = Leσ′, t = Leτ

′, where L is an arbitrary length scale,

we get

F (σ, τ) =√N +

∫ σ

−∞dσ′∫ τ

−∞dτ ′(MNF (σ′, τ ′)P (B)(σ′, τ ′) +

√MNF (σ′, τ ′)P (F )(σ′, τ ′)

),

F (σ, τ) =√M +

∫ σ

−∞dσ′∫ τ

−∞dτ ′(√

MNF (σ′, τ ′)P (F )(σ′, τ ′) +MNF (σ′, τ ′)P (B)(σ′, τ ′)).

(4.8)

with

P (F )(σ, τ) = −(

2π

k

)Γ(1/2− ε)(µL)2ε

25/2−επ3/2−εcos θ/2

cosϕ/2(eτ∂σ + eσ∂τ )

e−12

(σ+τ)eε(σ+τ)

(cosh(σ − τ) + cosϕ)12−ε ,

P (B)(σ, τ) =

(2π

k

)2 Γ2(1/2− ε)(µL)4ε

23−2επ3−2ε

cos2 θ/2

cos2 ϕ/2

cos2 ϕ/2 e2ε(σ+τ)

(cosh(σ − τ) + cosϕ)1−2ε.

(4.9)

F and F obey the differential equations

∂σ∂τF (σ, τ) = MNF (σ, τ)P (B)(σ, τ) +√MNF (σ, τ)P (F )(σ, τ) ,

∂σ∂τ F (σ, τ) =√MNF (σ, τ)P (F )(σ, τ) +MNF (σ, τ)P (B)(σ, τ) ,

(4.10)

with boundary conditions F (−∞, τ)=F (σ,−∞)=√N and F (−∞, τ)= F (σ,−∞)=

√M .

Then we write x = σ − τ and y = (σ + τ)/2 and obtain

(1

4∂2y − ∂2

x

)F (x, y) = MNF (x, y)P (B)(x, y) +

√MNF (x, y)P (F )(x, y) ,

(1

4∂2y − ∂2

x

)F (x, y) =

√MNF (x, y)P (F )(x, y) +MNF (x, y)P (B)(x, y) ,

(4.11)

– 12 –

with

P (F )(x, y) =

(2π

k

)Γ(12 − ε

)(µL)2ε

(2π)3/2−εcos θ2cos ϕ2

e2εy{d

dx

[sinhx/2

(coshx+ cosϕ)12−ε

]− ε coshx/2

(coshx+ cosϕ)12−ε

},

P (B)(x, y) =

(2π

k

)2 Γ2(12 − ε

)(µL)4ε

(2π)3−2εcos2 θ2cos2 ϕ2

e4εy{

(coshx+ cosϕ)2ε

2− sinh2 x/2

(coshx+ cosϕ)1−2ε

}.

(4.12)

4.1 General solution in d = 3

For ε = 0 equations (4.11) can be decoupled easily since the kernels (4.12) are independent

of y. Indeed, by introducing

H(x, y) = F (x, y) + F (x, y), K(x, y) = F (x, y)− F (x, y), (4.13)

(4.11) are equivalent to

(1

4∂2y − ∂2

x

)H(x, y) =

(aW ′(x)− a2W 2(x) +

a2

2

)H(x, y) ,

(1

4∂2y − ∂2

x

)K(x, y) =

(−aW ′(x)− a2W 2(x) +

a2

2

)K(x, y).

(4.14)

with

a =

(2π

k

) √MN

23/2π

cos θ/2

cosϕ/2, (4.15)

and

W (x) =sinhx/2

(coshx+ cosϕ)12

. (4.16)

These equations can be solved using the separation variable method. Setting

H(x, y) = h(y)ψ+(x) , K(x, y) = k(y)ψ−(x) , (4.17)

we get

∂2yh(y) = 4(−E +

a2

2)h(y) ,

∂2yk(y) = 4(−E +

a2

2)k(y) ,

(4.18)

and(−∂2

x + a2W 2(x)− aW ′(x))ψ+(x) = Eψ+(x) ,

(−∂2

x + a2W 2(x) + aW ′(x))ψ−(x) = Eψ−(x) .

(4.19)

The solution of the y dependent equations is simply

h(y) = C1e2

√−E+a2

2y + C2e

−2

√−E+a2

2y ,

k(y) = C3e2

√−E+a2

2y + C4e

−2

√−E+a2

2y ,

(4.20)

– 13 –

with C1,2,3,4 constants which have to be fixed by imposing the boundary conditions as we

will discuss in the following. The x−dependent equations (4.19) can be seen as the two

Schroedinger equations of a supersymmetric quantum mechanical system [60], therefore

E and E are non-negative. In principle one could solve for these equations. However we

are only interested to consider the case in which the edges of the cusp extend to infinity,

i.e. S and T very large. In this limit x ∼ 0 and y is very large, thus we make the

ansatz E = E = 0 4 and we set ψ+(0) = ψ−(0) = 1 since they can be reabsorbed in the

normalization constants C1, C2, C3, C4.

Using (4.17), (4.20) and (4.13) we get

F (0, y) =C1 + C3

2e√

2ay +C2 + C4

2e−√

2ay ,

F (0, y) =C1 − C3

2e√

2ay +C2 − C4

2e−√

2ay.

(4.21)

We fix the constants C1, . . . , C4 by matching the perturbative result. In d = 3 there are

UV divergences coming from the integration regions close to the cusps. To isolate this

divergence we set smin = tmin = δ which means ymin = ln δL ≡ −L0. At tree level,

obviously

F (0)(0,−L0) =√N, and F (0)(0,−L0) =

√M. (4.22)

Inserting these conditions in (4.21) we obtain

F (0, y) =√Ne√

2a(y+L0) − C2 + C4

2sinh√

2a(y + L0)e√

2aL0 ,

F (0, y) =√Me√

2a(y+L0) − C2 − C4

2sinh√

2a(y + L0)e√

2aL0 .

(4.23)

The two remaining constants are determined by matching the first order in the coupling

(which is contained in a) of our solution (4.23) with the first iteration of the Bethe-Salpeter

equations (4.8) at the same order

F (1)(τ, τ) = M√N

∫ τ

−L0

dσ′∫ τ

−L0

dτ ′PF (σ′, τ ′) ,

F (1)(τ, τ) = N√M

∫ τ

−L0

dσ′∫ τ

−L0

dτ ′PF (σ′, τ ′) ,(4.24)

which gives

C2 =

√N +

√M

2(1−A) ,

C4 =

√N −

√M

2(1 +A) ,

(4.25)

where

A = limy→∞L0→∞

√MN√

2a(y + L0)

∫ y

−L0

dσ′∫ y

−L0

dτ ′PF (σ′, τ ′) . (4.26)

4In the weak coupling limit for positive energy values the solutions of (4.20) become oscillatory.

– 14 –

In the appendix B we compute this integral and we find A = 1, thus C2 = 0 and C4 =√N −

√M . Inserting this result in (4.23) we finally obtain

〈W+〉 =

√NF +

√MF

N +M= cosh

√2a(y + L0) +

2√MN

N +Msinh√

2a(y + L0) .

〈W−〉 =

√NF −

√MF

N −M = cosh√

2a(y + L0) .

(4.27)

In order to extract the cusp anomalous dimension at arbitrary ϕ from the result (4.27),

we have simply to recast them in a suitable form to single out the logarithmic divergence.

Going back to the original (dimensionful) variables, we define

T = S = Λ−1IR = Ley → y = − logLΛIR = log

T

L;

δ = Λ−1UV = Le−L0 → L0 = logLΛUV = log

L

δ,

(4.28)

where ΛIR is the natural IR cut-off (associated to the length of the cusp) and ΛUV = 1δ is the

UV cut-off (cutting-off the cusp, where ladders collapse). The length scale L, introduced

previously for dimensional reason, will not play any role in the following. With the above

definitions we get

(y + L0) = logT

δ= log

ΛUVΛIR

. (4.29)

We can finally rewrite the expectation value of the cusped Wilson loop in the suggestive

form as follows:

〈W+〉 =(√M +

√N)2

2(M +N)e√

2a logΛUVΛIR +

(√M −

√N)2

2(M +N)e−√

2a logΛUVΛIR ,

〈W−〉 =1

2e√

2a logΛUVΛIR +

1

2e−√

2a logΛUVΛIR .

(4.30)

We have exactly reproduced the double-exponential structure found at two-loop in [57]: it

comes from an all-order computation, in a particular limit, strongly supporting the mixing

picture. In Section 5 we will discuss how the cusp anomalous dimension is related to this

result.

4.2 The straight-line limit: the solution for ε 6= 0

We consider here the straight-line limit ϕ = 0: remarkably the system enjoys supersym-

metry for any value of ε and the cusp anomalous dimension can be computed exactly. To

show this fact we first perform the change of variable x → ix′ and y → 12y′. Then using

(4.13), eqs. (4.11) become

�H(x′, y′) =[~∇W (x′, y′) · ~∇W (x′, y′)−�W (x′, y′)

]H(x′, y′) ,

�K(x′, y′) =[~∇W (x′, y′) · ~∇W (x′, y′) + �W (x′, y′)

]K(x′, y′) ,

(4.31)

– 15 –

with ~∇ = (∂x′ , ∂y′), � = ∂2x′ + ∂2

y′ and

W (x′, y′) =2ε−1/2aε

εeεy′cos2ε x

′

2, aε =

(2π

k

)√MN

Γ(1/2− ε)(µL)2ε

(2π)3/2−ε cos θ/2 . (4.32)

These equations are the Schroedinger equations of the two bosonic sectors of a two-

dimensional N = 2 supersymmetric quantum mechanics. The wave function of the ground

state can be exactly found and gives

H(x, y) =C1e−W (−ix′,2y′) = C1e

− 2ε−1/2aεε

e2εy cosh2ε x2 ,

K(x, y) =C2eW (−ix′,2y′) = C2e

2ε−1/2aεε

e2εy cosh2ε x2 ,

(4.33)

with C1 and C2 normalization constants. Thus, using (4.13), one finds

F (x, y) =C1

2e−

2ε−1/2aεε

e2εy cosh2ε x2 +

C2

2e

2ε−1/2aεε

e2εy cosh2ε x2 ,

F (x, y) =C1

2e−

2ε−1/2aεε

e2εy cosh2ε x2 − C2

2e

2ε−1/2aεε

e2εy cosh2ε x2 .

(4.34)

Here too we use the boundary conditions

F (0,−∞) =√N , and F (0,−∞) =

√M , (4.35)

to fix the constants in (4.34), getting

C1 =√N +

√M ,

C2 =√N −

√M .

(4.36)

The traced and supertraced operators for any x and y are

〈Wϕ=0+ 〉 =

√NF +

√MF

N +M= coshVε(x, y) +

2√MN

M +NsinhVε(x, y) ,

〈Wϕ=0− 〉 =

√NF −

√MF

N −M = coshVε(x, y) ,

(4.37)

where

Vε(x, y) = −W (−ix, 2y) = −2ε−1/2aεε

e2εy cosh2ε x

2. (4.38)

In order to consider the infinite cusped Wilson loop, recalling the original variables of the

Bethe-Salpeter integrals, we have to set x, y = 0, then we have

〈Wϕ=0+ 〉 =

(√M +

√N)2

2(M +N)e−

2ε−1/2

εaε +

(√M −

√N)2

2(M +N)e

2ε−1/2

εaε ,

〈Wϕ=0− 〉 =

1

2e−

2ε−1/2

εaε +

1

2e

2ε−1/2

εaε .

(4.39)

– 16 –

5 The determination of Γcusp(ϕ)

Above we have seen that the quantum expectation value of our cusped Wilson loop or-

ganizes itself as a double exponential in the limit (1.2) (see (4.30) or (4.39)). Here we

want to extract the cusp anomalous dimension from the Bethe-Salpeter results. We find

convenient to express the traced and supertraced operators WBa , a = ±, on a basis WB

i ,

i = 1, 2, whose elements renormalize multiplicatively, i.e.

WBi =Z(i)

cusp WRi . (5.1)

The two sets of bare operators are related by a linear transformation depending only on

the ranks of the gauge groups

WBa =AaiWB

i . (5.2)

From the explicit solution (4.30) or (4.39) of the Bethe-Salpeter equation and (3.5) one

reads

A =1

2(M +N)

(√

M +√N)2 (√

M −√N)2

M +N M +N

, (5.3)

while the regulator dependent parts enter in the Z(i)cusp (in this discussion we can neglect

the additional divergences subtracted by Zopen since they are suppressed in the limit of

large imaginary θ).

In the ϕ = 0 case one has

Z(1)cusp = e−

2ε−1/2

εaε , Z(2)

cusp = e2ε−1/2

εaε , (5.4)

thus

Γ(1)cusp =µ

∂

∂µlog(Z(1)

cusp) = −√MN

κcos

θ

2= −

√λ1λ2 ,

Γ(2)cusp =µ

∂

∂µlog(Z(2)

cusp) =

√MN

κcos

θ

2=

√λ1λ2 ,

(5.5)

where the definition (4.32) for aε has been used and λ1 and λ2 are the two effective coupling

constants introduced in (1.2). In the case of ABJM, where N = M , the two coupling

constants of course coincide: λ1 = λ2 ≡ λ.

In order to extract the cusp anomalous dimension at arbitrary ϕ, we have to consider

the result (4.30). We first consider the ABJM case. As already announced in Section 3.1

we observe a drastic simplification of the expectation value of the traced operator with

the disappearance of one of the two exponentials. In other words the trace renormalizes

multiplicatively. Conversely the supertrace stills mixes with the trace. Therefore, for

M = N , we can associate a cusp anomalous dimension, Γcusp(ϕ, λ), directly to W+ though

the usual definition (3.1) and, using (4.15), we find:

〈W+〉ABJM = e√

2a logΛUVΛIR ⇒ Γcusp(ϕ) = −

√2a = −N

κ

cos θ/2

cosϕ/2= − λ

cos ϕ2. (5.6)

– 17 –

In the general situation, when N 6= M , the expectation value of both W± contains a

double exponential, and the coefficients of the divergent logarithms in the exponentials are

identified with Γ(1,2)cusp (ϕ):

Γ(1)cusp(ϕ) =−

√λ1λ2

cos ϕ2, Γ(2)

cusp(ϕ) =

√λ1λ2

cos ϕ2, (5.7)

as explained in more details in the case ϕ = 0. Moreover the result (5.7) is perfectly

consistent with the limit (5.5) and for ϕ = θ = i∞, Γ(1,2)cusp vanish as expected!

Some remarks on the results (4.30) and (4.39) and the consequent form of Γ(i)cusp are now

in order. First we analyze the structure of the exponentiation in the ABJ case: looking

at our explicit calculation, one could expect that the positive cusp anomalous dimension

dominates, while the negative one gives a subleading contribution. On the other hand they

appear on the same footing in our computations and, much more crucially, consistency

with perturbative results requires the presence of both of them. However we believe that

the simple relation Γ(1)cusp(ϕ) = −Γ

(2)cusp(ϕ) implied by (5.7) does not survive when the

subleading corrections in θ are included. A second important observation concerns the

actual functional form of Γcusp(ϕ). Let us concentrate for the moment on the N = M

case. The final expression (5.6) is just the exponentiation of the one-loop result, namely

the leading cusp divergence undergoes to an abelian exponentiation in the ladder limit.

This result is completely different from the analogous N = 4 SYM resummation, where

an highly non-trivial function appears at this order, even for ϕ = 0. The reason relies

of course in the supersymmetric structure of the effective Schroedinger equation but it

has also a perturbative explanation: fermionic and bosonic diagrams do not exponentiate

in an abelian way by themselves and it is their delicate balance that, order by order in

the coupling constant, generates this nice behavior. We have checked explicitly at three-

loop in perturbation theory this fact. The N 6= M situation presents instead a slightly

more involved structure: we have still an abelian-like exponentiation at this order, but

when expressed in terms of the two (scaled) ’t Hooft couplings λ1, λ2 it appears through

a square root of their product. This is a further effect of the diagonalization process and

at moment we do not have a satisfying explanation from general principles. We stress that

from the point of view of the original CS level k the exponentiation is one-loop as well. A

third and, may be, more interesting remark is related to the strong-coupling limit and the

connection with string theory. To be concrete, we shall consider the simpler case of ABJM:

because of the abelian-like exponentiation we do not have any non-trivial interpolation

between weak and strong-coupling and the scaling limit does not match the√λ behavior

of string theory. At variance with N = 4 SYM the scaling limit does not seem to commute

with the strong-coupling limit, a fact that in four-dimensions was not expected a priori

(see the comments in the original computation [39]). We hope to come back soon on this

point when the subleading contributions will be computed.

– 18 –

6 Conclusions

In this paper we have studied a cusped Wilson loop in N = 6 Super Chern-Simons theory,

constructed with lines that are 1/2 BPS. We have computed the associated cusp anomalous

dimension in a scaling limit in which ladder diagrams dominate: because of the 1/2 BPS

character of the two halves, we have both bosonic and fermionic ladder exchanges and their

resummation is encoded into a coupled Bethe-Salpeter equation. We have seen that it can

be mapped into a supersymmetric Schroedinger equation whose ground state solution pro-

vide an exact expression for the cusp anomalous dimensions. Actually we found that, in the

general N 6= M case, the traced Wilson loop undergoes through a double-exponentiation,

as first observed in [57]. This has been interpreted as an operator mixing under cusp

renormalization: we have associated to the eigenvalues of the mixing matrix two indepen-

dent cusp anomalous dimensions. The final result is very simple and the exponentiations

are abelian, the cusp anomalous dimensions are one-loop exact up diagonalization. The

strong-coupling limit is therefore trivial and we do not find consistency with string theory

computation [61]: we argue that the scaling limit considered here does not commute with

the strong-coupling limit. Concerning abelian exponentiation, a similar phenomenon has

been observed recently [62] in studying N = 4 SYM cusped Wilson loops in k-symmetric

representations: at large N and k planar diagrams dominate and the exponentiation is of

abelian type.

The obvious follow-up of the present work is to take into account the subleading corrections

to the scaling limit: in [58] a systematic approach to this computation has been developed

in the N = 4 SYM case and it should be possible to perform an analogous investigation

here. Preliminary results seem promising. It would be interesting to see if the supersym-

metric structure we have found is preserved beyond leading order: in any case we expect a

non-trivial modification of the relation Γ(1)cusp(ϕ) = −Γ

(2)cusp(ϕ). Another direction consists

in checking the exponential structure at three-loop: the mixing we have observed here

prescribes an exponentiation with definite group-dependent coefficients (see eq. (4.30)),

that appear to be the same both in the scaling limit and in the general two-loop result

[57]. It would be of course nice to have a deeper understanding for the occurrence of the

mixing coefficients: a closer look at the supersymmetric quantum mechanics discussed in

[8], where the 1/2 BPS line is obtained from a Higgsing procedure, should be probably

useful for this task.

More ambitiously, one would like to approach the generalized cusp anomalous dimension

in ABJ(M) theory from a general point of view, with the hope to obtain other all-order

result by integrability or localization. In four-dimensions a particularly powerful approach

has been pushed forward recently [63, 64], applying to cusped Wilson loop the technique of

the quantum spectral curve. Beautiful results have been obtained for the Bremsstrahlung

function and the quark-anti-quark potential. It would be nice to extend this approach to

ABJ(M) case, in which the quantum spectral curve has been already studied [65]. It should

be also possible to extend the TBA equations derived in [28, 29] in the three-dimensional

– 19 –

context, taking advantage of the investigations presented in [66–68].

Acknowledgements

We thank Daniele Marmiroli for participating to the early stages of this work and Grisha

Korchemsky for a very useful and enlightening discussion at GGI. These investigations

have been supported in part by INFN and COST Action MP1210 ”The String Theory

Universe”.

– 20 –

A N = 6 three-dimensional Chern-Simons-matter theory

U(N) U(M)A

ICI

ICI

A

_

_

Figure 4: Quiver diagram for ABJ(M) theory.

The field content of the ABJ(M) theory can be schematically represented in the quiver

in Figure 4. The gauge sector consists of two gauge fields (Aµ)ij and (Aµ)i

jbelonging

respectively to the adjoint of U(N) and U(M). We denote by i, i the gauge indices in

the fundamental of the first and the second gauge group respectively. The matter sector

instead contains the complex fields (CI)ij and (CI )i

jas well as the fermions (ψI )i

j and

(ψI)ij. The fields (C, ψ) transform in the (N, M) of the gauge group U(N)×U(M) while

the pair (C, ψ) lives in the (N,M). The additional capitol index I = 1, 2, 3, 4 belongs to

the R-symmetry group SU(4). The ABJ(M) action is

SABJ(M) = SCS + Sgf + SMatter + SFint + SBint , (A.1)

where

SCS =− i κ4π

∫d3xεµνρ

[Tr(Aµ∂νAρ + 2

3 iAµAνAρ)− Tr(Aµ∂νAρ + 23 iAµAνAρ)

],

Sgf =κ

4π

∫d3x

[1ξTr (∂µAµ)2 + Tr (∂µcDµc)− 1

ξTr(∂µAµ)2 + Tr(∂µ¯cDµc

)],

SMatter =

∫d3x

[Tr(DµCID

µCI)

+ iTr(ψI /DψI

)],

(A.2)

and

SFint =− 2πi

κ

∫d3x

[Tr(CICIψJ ψ

J)− Tr(CICI ψJψJ) + 2Tr(CIC

J ψIψJ)

−2Tr(CICJψI ψJ)− εIJKLTr(CI ψJ CKψL) + εIJKLTr(CIψJCKψL)

],

SBint =− 4π2

3κ2

∫d3x

[Tr(CIC

ICJ CJCKC

K) + Tr(CICICJCJ C

KCK)

+4Tr(CICJCKC

ICJ CK)− 6Tr(CIC

JCJ CICKC

K)],

(A.3)

where ε1234 = ε1234 = 1 and κ is the Chern-Simons level. The matter covariant derivatives

are defined as

DµCI =∂µCI + i(AµCI − CIAµ), DµCI = ∂µC

I − i(CIAµ − AµCI),DµψI =∂µψI + i(AµψI − ψIAµ), Dµψ

I = ∂µψI − i(ψIAµ −AµψI).

(A.4)

– 21 –

Propagators and bilinears

The position-space propagators are obtained from those in momentum space (see e.g. [4])

by means of the following master integral

∫d3−2εp

(2π)3−2ε

eip·x

(p2)s=

Γ(

32 − s− ε

)

4sπ32−εΓ(s)

1

(x2)32−s−ε . (A.5)

In the Landau gauge, we have the following propagators

〈(Aµ) ji (x)(Aν) l

k (y)〉 =δliδjk

(2πi

κ

)εµνρ∂

ρxD(x− y),

〈(Aµ) j

i(x)(Aν) l

k(y)〉 =− δ l

iδjk

(2πi

κ

)εµνρ∂

ρxD(x− y).

〈(CI) ji (x)(CJ) l

k(y)〉 =δJI δ

li δ

j

kD(x− y),

〈(ψI) j

i(x)(ψJ) l

k (y)〉 =δJI δliδjkiγ

µ∂µD(x− y) ,

(A.6)

where

D(x− y) ≡ Γ(

12 − ε

)

4π32−ε

1

((x− y)2)12−ε . (A.7)

Computing the fermionic diagram contributing to the Wilson loop we have to deal with

the bilinear ηγη. Its expression in terms of the position along the line is [57]

(η2γµη1) =− 2

(η1η2)

[x1µ

|x1|+x2µ

|x2|− i x2

λ

|x2|x1ν

|x1|ε µλν

], (A.8)

where 1 and 2 denote two different points of the contour.

For our specific circuit all the possible products between η’s are the following

η1η1 = η2η2 = 2i, η1η2 = η2η1 = 2i cosϕ

2,

η2η1 = −2i sinϕ

2, η1η2 = 2i sin

ϕ

2. (A.9)

Here the indices 1 and 2 label the two different edges of the cusp.

B Useful integral

In the following we want to compute the quantity A appearing in the Section 4.1 and

defined by

A = limy→∞L0→∞

√MN√

2a(y + L0)

∫ y

−L0

dτ ′∫ y

−L0

dσ′P (F )(σ′, τ ′). (B.1)

– 22 –

It is convenient to have different upper bounds in the integrals. Using the definition ofP (F ) given by (4.9), we have:

∫ τ

−L0

dτ ′∫ σ

−L0

dσ′P (F )(σ′, τ ′) = −cos θ/2√2κ

1

cosϕ/2

∫ τ

−L0

dτ ′∫ σ

−L0

dσ′d

dσ′

(e

12 (τ

′−σ′)

(cosh(σ′ − τ ′) + cosϕ)12

)

= −cos θ/2√2κ

1

cosϕ/2

(I(σ, τ)− I(−L0, τ)

),

(B.2)

where we have computed the first integral using the total derivative and where we have

defined:

I(σ, τ) ≡∫ τ

−L0

dτ ′e

12

(τ ′−σ)

(cosh(σ − τ ′) + cosϕ)12

. (B.3)

We perform the change of variable z = eσ−τ′

and solve the first integral:

I(σ, τ) =√

2

∫ τ

−L0

dτ ′1

[e2(σ−τ ′) + 2e(σ−τ ′) cosϕ+ 1

]1/2

=−√

2

∫ e(σ−τ)

e(σ+L0)

dz

z

1

[z2 + 2z cosϕ+ 1]1/2

=−√

2

[− log

(1 + z cosϕ+

√z2 + 2z cosϕ+ 1

z

)]e(σ−τ)

e(σ+L0)

.

(B.4)

The second contribution is:

I(−L0, τ) = −√

2

[− log

(1 + z cosϕ+

√z2 + 2z cosϕ+ 1

z

)]e−(τ+L0)

1

. (B.5)

Summing up, we obtain:

I(σ, τ)− I(−L0, τ) =√

2

[G(σ − τ)−G(σ + L0)−G(−τ − L0) +G(0)

], (B.6)

where

G(x) = log(

1 + ex cosϕ+√e2x + 2ex cosϕ+ 1

). (B.7)

Now setting τ = σ = y, we have

I(y, y)− I(−L0, y) =√

2

[2G(0)−G(−y − L0)−G(y + L0)

]. (B.8)

For large y and L0, we can write the following expansion:

I(y, y)− I(−L0, y) ' −√

2(y + L0) + const +O(e−(y+L0)) . (B.9)

Therefore

√MN

∫ y

−L0

dτ ′∫ y

−L0

dσ′P (F )(σ′, τ ′) ' cos θ/2

cosϕ/2

√MN

κ(y + L0) =

√2a(y + L) . (B.10)

Recalling the definition (B.1), we obtain:

A = 1 . (B.11)

– 23 –

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