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Ladder operators can provide us deep insights into the characteristics of a system. In quantum mechanics, ladder operators relate families of solutions of the Schrödinger equation without requiring detailed knowledge of the solutions. When investigating the massive Klein-Gordon equation in curved spacetime, Cardoso et al. constructed generalized ladder operators for massive Klein-Gordon scalar fields in spacetimes with the conformal symmetry and showed that these operators can map a solution of the massive Klein-Gordon equation into that with different mass squared values [1, 2]. Thus, these operators are called mass ladder operators. In Ref. [3], Mück used a scalar curvature term to replace the mass squared term and built generalized ladder operators for the Klein-Gordon equation with a scalar curvature term. More recently, Katagiri and Kimura used ladder operators to investigate the quasinormal modes (QNMs) of a massive Klein-Gordon field in Bañados-Teitelboim-Zanelli (BTZ) spacetime, demonstrating that ladder operators can change the indices of QNM overtones under different boundary conditions [4].
The aforementioned studies on ladder operators focused only on Einstein gravity. It is of significant interest to generalize the investigation of ladder operators to Einstein-bumblebee gravity as it is a simple and effective theory involving Lorentz violation [5]. In the bumblebee model, Lorentz violation results from the dynamics of the bumblebee field, i.e., a single vector or axial-vector field
Bμ . In 2018, Casana and Cavalcante obtained an exact Schwarzschild-like black hole solution for this Einstein-bumblebee gravity and investigated some classical tests, including the advance of the perihelion, bending of light, and Shapiro's time-delay [6]. In 2020, Ding et al. obtained an exact Kerr-like solution and investigated the shadow of this Kerr-like black hole [7]. For a high dimensional spacetime, the authors of Ref. [8] obtained an exact AdS-like black hole solution and investigated the thermodynamics and phase transitions of this high dimensional Schwarzschild-AdS-like black hole. For three-dimensional spacetime, Ding et al. obtained an exact rotating BTZ-like black hole solution in the Einstein-bumblebee gravity theory and the central charges of the dual conformal field theory (CFT) on the boundary using the thermodynamic method [9]. Subsequently, Chen et al. investigated the quasinormal modes of a scalar perturbation around this rotating BTZ-like black hole in Einstein-bumblebee gravity and observed that the parameter breaking Lorentz symmetry imprints only in the imaginary parts of the quasinormal frequencies [10]. Hence, Einstein-bumblebee gravity has been explored extensively on various aspects, including exact solutions [11−19], geodesics and gravitational lensing [20−24], black hole shadows [25−29], quasinormal modes and stability [30−37], and energy extraction [38].In this study, we construct the mass ladder operators for the static BTZ-like black hole obtained in Ref. [9] in Einstein-bumblebee gravity, and we probe the QNM frequencies of the mapped modes using mass ladder operators for the scalar perturbation under Dirichlet and Neumann boundary conditions. We observe that the mass ladder operators, which depend on the Lorentz symmetry breaking parameter s in Einstein-bumblebee gravity, can map a solution of the massive Klein-Gordon equation for the BTZ-like black hole into that with different mass squared values and change indices of QNM overtones. Particularly, we find that a threshold value
sc exists for all the modes, i.e., the quasinormal frequency has an imaginary part only if the parameters<sc .The remainder of this article is organized as follows. In Sec. II, we first review the static BTZ-like black hole in Einstein-bumblebee gravity; we then calculate the QNMs and frequencies for a massive perturbation of the static BTZ-like black hole under the Dirichlet and Neumann boundary conditions. In Sec. III, we construct the mass ladder operators in the static BTZ-like spacetime, discuss the relationship between the mass ladder operators and QNMs, and analyze the effect of the overtone number n and Lorentz symmetry breaking parameter s on the mapped modes. Finally, in Sec. IV, we summarize our study.
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Let us begin by briefly reviewing the static BTZ-like spacetime in Einstein-bumblebee gravity obtained in [9]. The action in this theoretical model with a negative cosmological constant
Λ=−1l2 is given by [39−41]S=∫d3x√−g[R−2Λ2κ+ϱ2κBμBνRμν−14BμνBμν−V(BμBμ±b2)+LM].
(1) Here, R describes the Ricci scalar, and κ is a constant that has a relationship with the three-dimensional Newton's constant G, given by
κ=8πG .Bμν=∂μBν−∂νBμ is the strength of the bumblebee fieldBμ , andϱ is a coupling constant with a dimension ofM−1 in this model. In addition, we note that the potential V is minimum atBμBμ±b2=0 , where b is a real positive constant and the signs± determine whether the fieldbμ is timelike or spacelike. This minimum yields a nonzero vacuum value⟨Bμ⟩=bμ withbμbμ=∓b2 , which results in the breaking ofU(1) symmetry. Note that the nonzero vector backgroundbμ results in the violation of Lorentz symmetry [5, 39−41].From the action (1), we can obtain the static BTZ-like black hole solution in Einstein-bumblebee gravity [9]:
ds2=−f(r)dt2+1+sf(r)dr2+r2dφ2,
(2) with
f(r)=r2l2−M,
(3) where M denotes the mass of the black hole, and
s=ξb2 represents the spontaneous breaking of Lorentz symmetry owing to the Einstein-bumblebee vector field with the formbμ=(0,bξ,0,0) . Considering the determinant of the metricg=−(1+s)r2 , we find that the metric becomes degenerate ats=−1 . In the following analysis, we consider the constraints>−1 and compare the results with those fors=0 , i.e., without Lorentz symmetry breaking. The Hawking temperature of this BTZ-like black hole isTH=rH8πl2√1+s,
(4) with the horizon
rH=√Ml . For convenience, we will scalel=1 in the numerical calculation.In the static BTZ-like spacetime, the massive scalar field evolves according to
(∇μ∇μ−μ2)Ψ=0,
(5) where μ is the mass of the scalar field. Assuming that the massive scalar perturbation Ψ has the form
Ψ=e−iωt+imφψ(r),
(6) we can obtain the radial equation
d2ψ(r)dr2+[1r+f′(r)f(r)]dψ(r)dr+(1+s)[ω2f(r)2−m2r2f(r)−μ2f(r)]ψ(r)=0,
(7) which can be solved in terms of the hypergeometric functions [42]. We introduce a variable
z=(r2−r2H)/r2 ; thus, the radial Eq. (7) can be rewritten asz(1−z)d2ψ(z)dz2+(1−z)dψ(z)dz+(Az−B1−z−C)ψ(z)=0,
(8) where A, B, and C have the forms
A=ω2l4(1+s)4r2H,B=μ2l2(1+s)4,C=m2l2(1+s)4r2H.
(9) Rewriting the radial function
ψ(z) as the formψ(z)=zil2√1+s2rHω(1−z)1+√1+μ2l2(1+s)2F(z),
(10) we find that the function
F(z) satisfies the standard hypergeometric equationz(1−z)d2F(z)dz2+[c−(1+a+b)z]dF(z)dz+abF(z)=0,
(11) with
a=1+√1+μ2l2(1+s)2−il√1+s2rH(ωl−m),
(12) b=1+√1+μ2l2(1+s)2−il√1+s2rH(ωl+m),
(13) c=1−il2√1+srHω,
(14) where the scalar field mass must obey
μ2⩾μ2BF with the Breitenlohner-Freedman (BF) boundμ2BF=−1l2(1+s) . Thus, the general solution of the radial Eq. (8) can be given by a linear combination of hypergeometric functions F, i.e.,ψ(z)=zil2√1+s2rHω(1−z)1+√1+μ2l2(1+s)2[AF(a,b,c;z)+Bziωl2√1+srHF(a−c+1,b−c+1,2−c;z)],
(15) where A and B are two constants of integration. By imposing the ingoing-wave condition at the black hole horizon, we obtain
B=0 ; thus, the solution (15) has a more simple form:ψ(z)=Azil2√1+s2rHω(1−z)1+√1+μ2l2(1+s)2F(a,b,c;z).
(16) To probe the asymptotic behavior of
ψ(z) at infinity asr→∞ (i.e.,z→1 ), we can expand the hypergeometric functions at1−z [42], i.e.,lF(a,b,c;z)=Γ(c)Γ(c−a−b)Γ(c−a)Γ(c−b)F(a,b,a+b−c+1;1−z)+(1−z)c−a−bΓ(c)Γ(a+b−c)Γ(a)Γ(b)F(c−a,c−b,c−a−b+1;1−z).
(17) Thus, the general solution (16) can be expressed as
ψ(r)=AIψ(D)(r)+AIIψ(N)(r),
(18) where the constants
AI andAII areAI=AΓ(c)Γ(c−a−b)Γ(c−a)Γ(c−b),AII=AΓ(c)Γ(a+b−c)Γ(a)Γ(b).
(19) The functions
ψ(D)(r) andψ(N)(r) denote the forms of solution (16) that satisfy the Dirichlet and Neumann boundary conditions, respectively, at infinity [4]. The forms ofψ(D)(r) andψ(N)(r) areψ(D)(r)=(1−r2Hr2)il2√1+s2rHω(rHr)1+√1+μ2l2(1+s)×F(a,b,a+b−c+1;r2Hr2),
(20) ψ(N)(r)=(1−r2Hr2)il2√1+s2rHω(rHr)1−√1+μ2l2(1+s)×F(c−a,c−b,c−a−b+1;r2Hr2).
(21) With the Dirichlet boundary condition, we find
a=−n orb=−n and then obtain the quasinormal frequencies for the massive scalar perturbationω(D)=±ml−i2rHl2√1+s[n+12+12√1+μ2l2(1+s)].
(22) The signs
+ and− denote the left-moving and right-moving modes, respectively. As s increases, the absolute value of the imaginary part ofω(D) decreases for different n values [10], which is also shown in Fig. 1(a). Moreover, we also note that all the imaginary parts ofω(D) are negative, which means that the black hole is stable under the scalar perturbation with the Dirichlet boundary condition. Additionally, the corresponding wave-functionΨ(D) at infinity has the asymptotic behaviorFigure 1. (color online) Relationship between the imaginary part of the QNM frequencies and the Lorentz symmetry breaking parameter s with different overtone numbers n under different boundary conditions, where
rH=1 andμ2=−0.75 .Ψ(D)|r→∞=AI(rHr)1+√1+μ2l2(1+s)×[1+O(1/r2)]e−iω(D)t+imφ.
(23) Similarly, for the Neumann boundary condition, we obtain
c−a=−n orc−b=−n ; thus, the quasinormal frequencies for the scalar field areω(N)=±ml−i2rHl2√1+s[n+12−12√1+μ2l2(1+s)].
(24) As shown in Fig. 1(b), similar to the case of the Dirichlet boundary condition, the imaginary part of
ω(N) increases for different n values as s increases under the Neumann boundary condition, except for the fundamental moden=0 , i.e., the scalar field without nodes, where the imaginary part ofω(N) decreases slowly as s increases. Note that the asymptotic behavior of the wave-functionΨ(N) at infinity becomesΨ(N)|r→∞=AII(rHr)1−√1+μ2l2(1+s)×[1+O(1/r2)]e−iω(N)t+imφ.
(25) -
For the d-dimensional spacetime
(M,gμν) with the symmetry under the conformal transformationg_{\mu\nu}\rightarrow g'_{\mu\nu} = e^{2Q}g_{\mu\nu}u\nu} andQ=1d∇σζσ , a conformal Killing vector (CKV)ζμ exists that satisfies the conformal Killing equation∇μζν+∇νζμ=2d∇σζσgμν.
(26) Because the closed conformal Killing vector (CCKV)
ζμ satisfies an extra condition∇[μζν]=0,
(27) it satisfies the reduced Killing equation
∇μζν=1d∇σζσgμν.
(28) Recent investigations show that a CCKV
ζμ is an eigenvector of Ricci tensor with a constant eigenvalue χ [1, 2], i.e.,Rμνζν=χ(d−1)ζμ.
(29) We can introduce a one-parameter family of operators
Dk:=Lζ−kd∇μζμ,
(30) where
Lζ denotes the Lie derivative with respect toζμ , and k is a parameter. The commutation relation between the operatorDk and the d'Alembertian is given by[∇μ∇μ,Dk]=χ(2k+d−2)Dk+2Q(∇μ∇μ+χk(k+d−1)).
(31) Acting the commutator on a scalar field ϕ, we can obtain
Dk−2[∇μ∇μ+χk(k+d−1)]ϕ=[∇μ∇μ+χ(k−1)(k+d−2)]Dkϕ.
(32) The operator
Dk maps a scalar field ϕ with mass squaredμ2=−χk(k+d−1) to another scalar fieldDkϕ with a new mass squaredμ′2=−χ(k−1)(k+d−2) ; therefore,Dk can be considered a mass ladder operator, which shifts parameter k tok−1 in the mass squared term. Thus, this type of operator is called the mass lowering operatorDk− . Similarly, we can introduce the mass raising operatorDk+ , which shifts parameter k tok+1 in the mass squared term.For the BTZ-like spacetime in Einstein-bumblebee gravity (2), we obtain four independent CCKVs
ζ0=erHl2√1+st(1√r2−rH2∂t−r√r2−rH2rHl2√1+s∂r),
(33) ζ1=e−rHl2√1+st(1√r2−rH2∂t+r√r2−rH2rHl2√1+s∂r),
(34) ζ2=erHl√1+sφ(r2−rH2rHl√1+s∂r+1r∂φ),
(35) ζ3=e−rHl√1+sφ(−r2−rH2rHl√1+s∂r+1r∂φ),
(36) and the corresponding four mass ladder operators
D0,k=erHl2√1+st(1√r2−rH2∂t−r√r2−rH2rHl2√1+s∂r+k√r2−rH2rHl2√1+s),
(37) D1,k=e−rHl2√1+st(1√r2−rH2∂t+r√r2−rH2rHl2√1+s∂r−k√r2−rH2rHl2√1+s),
(38) D2,k=erHl√1+sφ(r2−rH2rHl√1+s∂r+1r∂φ−krrHl√1+s),
(39) D3,k=e−rHl√1+sφ(−r2−rH2rHl√1+s∂r+1r∂φ+krrHl√1+s).
(40) The mass ladder operators depend on s in Einstein-bumblebee gravity. In an AdS
d -like spacetime, the BF-boundμ2BF for a scalar field can be expressed as [1, 4]μ2BF=(d−1)24χ,χ<0.
(41) For the BTZ-like black hole spacetime in Einstein-bumblebee gravity (2), we find that χ is exactly equal to the BF-bound
μ2BF=−1l2(1+s) , and then the commutation relation (31) with the d'Alembertian becomes[∇μ∇μ,Di,k]=−2k+1l2(1+s)Di,k+23(∇μζμi)[∇μ∇μ−k(k+2)l2(1+s)].
(42) This means that Eq. (32) can be further simplified as
Di,k−2[∇μ∇μ−k(k+2)l2(1+s)]Ψ=[∇μ∇μ−(k−1)(k+1)l2(1+s)]Di,kΨ.
(43) As in [1, 2],
Di,k maps a solution Ψ to the Klein-Gordon equation with the mass squaredμ2=k(k+2)l2(1+s) into another solutionDi,kΨ with the mass squaredμ2=(k−1)(k+1)l2(1+s) , namely, it shifts k tok−1 in the mass squared term.To require that the characteristic exponents with the mass
Δ±=12[1±√1+μ2l2(1+s)] in (23) and (25) are real, we focus only on the mass squaredμ2⩾μ2BF . As in [4], we can introduce a parameter ν, which obeysμ2l2(1+s)=ν(ν+2).
(44) This means that
ν=−1±√1+μ2l2(1+s) . The relationship betweenμ2l2(1+s) and ν for different values of s is presented in Fig. 2, which shows that the functionμ2l2(1+s) is symmetric aboutν=−1 . Thus, without loss of generality, we focus only on the right part of the parabola whereν=−1+√1+μ2l2(1+s) withν∈[−1,∞) . For a given parameter ν, two values of k satisfyk(k+2)=ν(ν+2) , i.e.,k+=−2−ν,k−=ν.
(45) The corresponding mass ladder operators
Di,k+ andDi,k− shift ν toν+1 andν−1 , respectively.Now, we can discuss the mapped solutions resulting from the mass ladder operators acting on the original solution Ψ. Because the operators
D2,k± andD3,k± are non-global operators, they are not globally smooth in the φ direction. Thus, they change the quantum number m of the scalar field. However, the operatorsD0,k± andD1,k± can directly affect the frequencies of the scalar perturbation. Therefore, as in [4], we focus only on the effects ofD0,k± andD1,k± on the general solution of the scalar field under different boundary conditions.The general solution (16) of the scalar field can be rewritten as
Ψ=A(1−r2Hr2)−il2√1+s2rHω(rHr)2+νF(a,b,c;1−r2Hr2)e−iωt+imφ.
(46) The asymptotic behavior of the general solution near the horizon is
Ψ|r≃rH=2−il2√1+s2rHωA(r−rHr)−il2√1+s2rHω×[1+O(r−rH)]e−iωt+imφ.
(47) Under the Dirichlet or Neumann boundary conditions at infinity, the above general solution has the asymptotic behavior
Ψ|r≃∞=AI(rHr)2+ν[1+O(1/r2)]e−iωt+imφ(DBC),
(48) Ψ|r≃∞=AII(rHr)−ν[1+O(1/r2)]e−iωt+imφ(NBC).
(49) Acting the mass ladder operators
D0,k± andD1,k± on the exact solution (46), we can obtain the asymptotic behaviors of the mapped solutions near the horizon:D0,k±Ψ|r≃rH=C0,k±(r−rHr)−il2√1+s2rH(ω+irHl2√1+s)[1+O(r−rH)]e−i(ω+irHl2√1+s)t+imφ,
(50) D1,k±Ψ|r≃rH=C1,k±(r−rHr)−il2√1+s2rH(ω−irHl2√1+s)[1+O(r−rH)]e−i(ω−irHl2√1+s)t+imφ,
(51) where
C0,k± andC1,k± are two constants:C0,k±=iArHl4(1+s)(ω+irHl2√1+s)2−1−il2√1+s2rH(ω+irHl2√1+s)×[r2H(k±(2+k±)−ν(2+ν))+l4(1+s)(ω+ml−ik±rHl2√1+s)(ω−ml−ik±rHl2√1+s)], (52) C1,k±=−21−il2√1+s2rH(ω−irHl2√1+s)iωArH.
(53) Under the Dirichlet boundary condition, the asymptotic behaviors of the mapped solutions acted by the mass ladder operators
D0,k± andD1,k± at infinity can be expressed asD0,k+Ψ(D)|r≃∞=C(D)0,k+(rHr)3+ν×[1+O(1/r2)]e−i(ω(D)+irHl2√1+s)t+imφ,
(54) D1,k+Ψ(D)|r≃∞=C(D)1,k+(rHr)3+ν×[1+O(1/r2)]e−i(ω(D)−irHl2√1+s)t+imφ,
(55) D0,k−Ψ(D)|r≃∞=C(D)0,k−(rHr)1+ν×[1+O(1/r2)]e−i(ω(D)+irHl2√1+s)t+imφ,
(56) D1,k−Ψ(D)|r≃∞=C(D)1,k−(rHr)1+ν×[1+O(1/r2)]e−i(ω(D)−irHl2√1+s)t+imφ,
(57) where
C(D)0,k+=−AI(2+ν)l2√1+s2r2H[ω(D)−ml+irHl2√1+s(2+ν)]×[ω(D)+ml+irHl2√1+s(2+ν)],
(58) C(D)1,k+=AI(2+ν)l2√1+s2r2H[ω(D)−ml−irHl2√1+s(2+ν)]×[ω(D)+ml−irHl2√1+s(2+ν)],
(59) C(D)0,k−=2(1+ν)l2√1+sAI,C(D)1,k−=−2(1+ν)l2√1+sAI.
(60) Combining with Eqs. (50), (51), and (54)−(57), we find that
D0,k± andD1,k± shift ω toω+−rHl2√1+s andω−−rHl2√1+s , respectively, while keeping m unchanged. Thus, the quasinormal frequencies of the mapped solutions of the scalar field by the mass ladder operatorsD0,k± andD1,k± areω(D)0,1=ω(D)±−rHl2√1+s,
(61) with
ω(D)=±ml−irHl2√1+s(2n+ν+2).
(62) In particular, when the mass ladder operator
D0,k+ acts on the fundamental mode, the mapped mode vanishes rather than becomes a “negative overtone mode” owing to the restriction imposed by (58) as in [4]. Figure 3 shows the change of quasinormal frequencies of the mapped modesD0,k±Ψ(D) andD1,k±Ψ(D) with the parameter s, which is similar to that of the original modesΨ(D) . Notice that, for all the modes withμ2=−0.75 andl=1 , a threshold valuesc=13 exists, corresponding to a vertical dashed line in each panel, which means that all the modes stop increasing their imaginary parts when they reach this vertical line. The threshold valuesc=−(1+1μ2l2) depends only on the mass of the scalar field and increases with the mass squared termμ2 , as shown in Fig. 4 forμ2⩽0 .Figure 3. (color online) Relationship between the imaginary part
Im [ω(D)] and s forΨ(D) ,D0,k±Ψ(D) , andD1,k±Ψ(D) for different values of n, whererH=1 andμ2=−0.75 .Figure 4. (color online) Relationship between
Im [ω(D)] and s forΨ(D) ,D0,k−Ψ(D) , andD1,k±Ψ(D) with different massesμ2 , wheren=0 andrH=1 .Under the Neumann boundary condition, we can obtain the asymptotic behaviors of the mapped solutions by mass ladder operators
D0,k± andD0,k± at infinity:D0,k+Ψ(N)|r≃∞=C(N)0,k+(rHr)−1−ν×[1+O(1/r2)]e−i(ω(N)+irHl2√1+s)t+imφ,
(63) D1,k+Ψ(N)|r≃∞=C(N)1,k+(rHr)−1−ν×[1+O(1/r2)]e−i(ω(N)−irHl2√1+s)t+imφ,
(64) D0,k−Ψ(N)|r≃∞=C(N)0,k−(rHr)1−ν×[1+O(1/r2)]e−i(ω(N)+irHl2√1+s)t+imφ,
(65) D1,k−Ψ(N)|r≃∞=C(N)1,k−(rHr)1−ν×[1+O(1/r2)]e−i(ω(N)−irHl2√1+s)t+imφ,
(66) where
C(N)0,k+=−2(1+ν)l2√1+sAII,C(N)1,k+=2(1+ν)l2√1+sAII,
(67) C(N)0,k−=AIIνl2√1+s2r2H[ω(N)−ml−irHl2√1+sν]×[ω(N)+ml−irHl2√1+sν],
(68) C(N)1,k−=−AIIνl2√1+s2r2H[ω(N)−ml+irHl2√1+sν]×[ω(N)+ml+irHl2√1+sν].
(69) Combining with Eqs. (50), (51), (63)−(66), we find that
D0,k± andD1,k± shift ω toω+irHl2√1+s andω−irHl2√1+s , respectiveley, while keeping m unchanged. Thus, the quasinormal frequencies of the mapped solutions of the scalar field by the mass ladder operatorsD0,k± andD1,k± areω(N)0,1=ω(N)±irHl2√1+s,
(70) with
ω(N)=±ml−irHl2√1+s(2n−ν).
(71) Similarly, under the Neumann boundary condition, no ''negative overtones'' are generated from the fundamental modes for
D0,k− as in the case with the Dirichlet boundary condition. For the fundamental mode withμ2<0 , the mass ladder operators do not change the BF-boundμ2BF . Furthermore, as in the Dirichlet boundary condition, we find that the same threshold valuesc=1/3 also exists for the mapped and original modes. The threshold valuesc for all the mapped modes with the same mass squared term is a constant. Thus, the mass ladder operators maintain the threshold valuesc for the mapped modes.Moreover, we observe that, for the
n=0 case in Fig. 5, the modesD0,k+Ψ(N) andΨ(N) decrease with the increase in s, which differ with the other modes. Therefore, we focus on then=0 case as shown in Fig. 6, which indicates that the threshold valuesc increases with the increase inμ2 . Note that the imaginary parts of the quasinormal frequencies of the mapped modes caused byD0,k+Ψ(N) are all positive, which indicates that all the fundamental modes ofD0,k+Ψ(N) are unstable. -
In this study, we investigated mass ladder operators constructed from the conformal symmetry of the static BTZ-like black hole in Einstein-bumblebee gravity. Additionally, we probed the QNM frequencies of the mapped modes using mass ladder operators for the scalar perturbation under the Dirichlet and Neumann boundary conditions. We found that the mass ladder operators, which can change not only the mass squared but also the QNM frequencies, depend on the Lorentz symmetry breaking parameter s in Einstein-bumblebee gravity. We observed that the imaginary parts of the frequencies shifting by mass ladder operators increase with an increase in s. The changes of quasinormal frequencies of the mapped modes with s are similar to those of the original modes under two boundary conditions. However, under the Neumann boundary condition, the imaginary parts of the quasinormal frequencies of the mapped modes caused by the mass ladder operator
D0,k+ are all positive, which implies that all the corresponding fundamental modes are unstable. Moreover, mass ladder operators do not change the BF-boundμ2BF as in the usual BTZ black hole, and the threshold value of the Lorentz symmetry breaking parametersc , where the modes stop increasing their imaginary parts, increases with an increase in mass squaredμ2 . These results can aid us in further understanding the conformal symmetry and Lorentz symmetry breaking of the static BTZ-like black hole in Einstein-bumblebee gravity.
Mass ladder operators and quasinormal modes of the static BTZ-like black hole in Einstein-bumblebee gravity
- Received Date: 2024-07-26
- Available Online: 2025-01-15
Abstract: We investigate mass ladder operators for the static BTZ-like black hole in Einstein-bumblebee gravity and probe the quasinormal frequencies of the mapped modes using mass ladder operators for a scalar perturbation under Dirichlet and Neumann boundary conditions. We find that the mass ladder operators depend on the Lorentz symmetry breaking parameter, and the imaginary parts of the frequencies shifted by the mass ladder operators increase with the increase in the Lorentz symmetry breaking parameter under the two boundary conditions. Note that, under the Neumann boundary condition, the mapped modes caused by the mass ladder operator