Dissipative preparation of distributed steady ... - OSA Publishing

Vol. 25, No. 1 | 9 Jan 2017 | OPTICS EXPRESS 88

Dissipative preparation of distributed steady entanglement: an approach of unilateral qubit driving Z HAO J IN , 1 S HI -L EI S U, 2 A I -D ONG Z HU, 3 H ONG -F U WANG , 3 AND S HOU Z HANG 1,3,* 1 1Department

of Physics, Harbin Institute of Technology, Harbin 150001, China of Physical Science Engineering and Key Laboratory of Materials Physics of Ministry of Education of China, Zhengzhou University, Zhengzhou 450052, China 3 Department of Physics, College of Science, Yanbian University, Yanji, Jilin 133002, China 2 School

* [email protected]

Abstract: We propose a nonlocal scheme for preparing a distributed steady-state entanglement of two atoms trapped in separate optical cavities coupled through an optical fiber based on the combined effect of the unitary dynamics and dissipative process. In this scheme, only the qubit of one node is driven by an external classical field, while the other one does not need to be manipulated by an external field. This is meaningful for long distance quantum information processing tasks, and the experimental implementation is greatly simplified due to the unilateral manipulation on one node and the process of entanglement distribution can be avoided. This guarantees the absolute security of long distance quantum information processing tasks and makes the scheme more robust than that based on the unitary dynamics. We introduce the purity to characterize the mixture degree of the target steady-state. The steady entanglement can be obtained independent of the initial state. Furthermore, based on the dissipative entanglement preparation scheme, we construct a quantum teleportation setup with multiple nodes as a practical application, and the numerical simulation demonstrates the scheme can be realized effectively under the current experimental conditions. c 2017 Optical Society of America OCIS codes: (270.0270) Quantum optics; (270.5565) Quantum communications; (270.5580) Quantum electrodynamics; (270.5585) Quantum information and processing.

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#280010 Journal © 2017

http://dx.doi.org/10.1364/OE.25.000088 Received 8 Nov 2016; revised 6 Dec 2016; accepted 10 Dec 2016; published 3 Jan 2017


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1.

Introduction

Quantum entanglement plays a crucial role in performing quantum information processing [1,2]. From a fundamental perspective, it is a nonclassical effect which is indispensable for understanding fundamental quantum physics. From a technological perspective, it is useful for enhanced measurement techniques and is the basic requirement for transferring quantum information between different nodes in a quantum network. The preparation and storage of the entanglement between distant quantum nodes leaves a great challenge due to environmental decoherence as well as the imperfection of quantum system itself [3]. Recently, the system of atom-cavity-fiber [4–6] has been attracted much attentions due to the application in long distance and large-scale quantum information processing, such as distributed quantum computation [7, 8], quantum entanglement preparation [9–14], and the quantum communication [15]. For these unitary-dynamics-based schemes, atomic spontaneous emission, cavity decay and the fiber loss are three main dissipative factors which would affect the practical efficiency of these schemes. Traditionally, dissipation is considered as a detrimental effect in quantum information processing, however, recent theories and experiments show an interesting fact that the dissipation can be used as resources for quantum computation and entanglement generation [16–47]. In contrast with the unitary-dynamics-based schemes, here the dissipation plays a positive role in the preparation process so that the entanglement is robust against decoherence. Furthermore, since the entangled state appears as a steady-state, it is unnecessary to introduce an additional unitary feedback mechanism to stabilize it. Another merit of this approach is that we do not require specifying the initial state and controlling the evolution time


accurately. Particularly, In 1999, Plenio et al. and Cabrillo et al. proposed schemes to prepare entanglement via dissipation [16, 17]. Immediately afterward, several schemes were suggested to study the entanglement in dissipative quantum system. In 2011, Kastoryano et al. proposed a dissipative scheme for preparing a maximally entangled state of two -type atoms trapped in one optical cavity [18], whose results are better than that based on the unitary dynamics. Subsequently, Shen et al. generalized this scheme to the coupled cavity system [32, 34] and atomcavity-fiber system [33]. In 2013, Reiter et al. [45] demonstrated that two transmon qubits can be driven into the steady Bell state by combining the effective two-photon process induced by microwave driving with the photon loss. In 2014, Zheng et al. proposed two schemes to prepare the maximal entanglement between two atoms coupled to a decaying resonator [35, 47]. All of the previous schemes need to use four or more classical fields to drive two qubits simultaneously, and relative amplitudes and phases of the driving fields applied to different qubits should be accurately set. Moveover, the experimental demonstrations of dissipative preparation of entanglement has also been reported in ion traps [44, 49], superconducting circuits [52], and the collective spin degrees of freedom of large atomic ensembles [48]. In this paper, we propose a feasible scheme to prepare and stabilize a maximally entangled Bell state in separate optical cavities coupled through an optical fiber, where only one qubit is needed to be driven by two classical fields with well-chosen frequencies. With currently achievable experiment parameters, the numerical simulation demonstrates that the distributed steadystate entanglement can be obtained with high fidelity, purity and CHSH correction. Compared with previous schemes [47], the present one does not require simultaneous driving to both qubits, which is a basic requirement to perform state transfer and quantum gate operation between separate nodes of a quantum network. Different from the Ref. [35], the present one generalize the idea that two atoms are trapped in a single cavity to the case when two atoms trapped into two separate cavities connected through an optical fiber. It is significant for quantum networks since only one unilateral driving on the qubit in one node is required with classical fields. The organization of the rest paper is as follows: In Sec. 2, we present the details of our model and show the master equation describing the dynamics of the open dissipative system in Lindblad form. In Sec. 3, we first expatiate the system’s dressed state subspace. Then, a steady Bell state is generated by using the classical laser fields to drive one qubit within the dressed space. Furthermore, we assess the performance of this scheme through numerically calculating the fidelity, purity and Clauser-Horne-Shimony-Holt (CHSH) correlation, respectively. In Sec. 4, we construct a quantum teleportation scheme with multiple nodes and calculate the variation of the fidelity of teleportation with the increasing of the node number n. Finally, we present the conclusions inferred from this paper in Sec. 5. 2.

Theoretical model

We consider a atom-cavity-fiber coupling system consisting of two distant cavities connected by a single-transverse-mode optical fiber, as shown in Fig. 1. Each atom has ground state |g and excited state |e with the corresponding energies 0 and ω0 , respectively. The atomic transition |g → |e is coupled resonantly to the cavity mode with the coupling constant g, and the first atom driven by two off-resonance optical lasers with corresponding detuning Δk (k = 1, 2). In the short fiber limit Lν/(2πc) ≤ 1, where L is the length of the fiber and ν is the decay rate of the cavity field into a continuum of the fiber modes, only one fiber mode essentially interacts with the cavity modes. For simplicity, we assume the interaction between cavity mode and fiber mode is resonant. Thus, in a rotating frame, the Hamiltonian of the system could be written as (setting = 1 throughout this paper) H = H0 + Hc, f + Ha,c + Hcl , H0 =

i=1,2

ω0 |ei ei | +

j= A, B

ωa a †j a j + ωb b† b,

(1)


γ

γ

Δ2

κ

ω1

g

e2

e1

Δ1

κ

g

ω2

β

g1

CavityA

g2

CavityB

Fig. 1. Experimental setup and level diagram of the atoms. Two atoms resonantly interact with quantized field, respectively. The first atom is driven by two classical fields. γ, κ, and β denote the atomic spontaneous emission rate, cavity decay rate and fiber loss rate, respectively.

Hc, f = ν(ba †A + ba †B ) + H.c.,

(2)

Ha,c = g(S1† a A + S2† a B ) + H.c.,

(3)

Hcl =

k=1,2

Ωk e −iω k t S1† + H.c.,

(4)

where S1† = |e1 g1 | and S2† = |e2 g2 |, |ei and |gi are the excited and ground states of the ith qubit, respectively. a j and a †j denote the annihilation and creation operators for the optical mode of cavity j, respectively. b and b† denote the annihilation and creation operators for the fiber mode, respectively. Ωk and ωk represent the amplitude and frequency of the kth driving field, respectively. ωa and ωb denote the frequencies of cavity mode and fiber mode, respectively. In order to investigate dynamics of the system further, we introduce the non-local bosonic modes √ 2 (a A − a B ), c= 2 √ 1 c1 = (a A + a B + 2b), 2 √ 1 c2 = (a A + a B − 2b), (5) 2 √ √ corresponding frequencies ωa , ωa + 2ν and ωa − 2ν. These modes are not coupled with each other, but interact with the atoms because of the contributions of the cavity fields. To simplify the dynamics of the system, In the interaction picture with respect to H0 + Hc, f , the Hamiltonian describing the atom-cavity interaction is √ √ √ 1 Ha,c = g(e −i 2νt c1 + e i 2νt c2 + 2c)S1† 2 √ √ √ 1 + g(e −i 2νt c1 + e i 2νt c2 − 2c)S2† + H.c., (6) 2 meanwhile, under the condition of |ν| g, the bosonic mode c is resonant with the two qubits, while the bosonic modes c1 and c2 is largely dispersive with the two qubits. Therefore, the interaction Hamiltonian of atom-cavity reduces to √ 2 g(S1† − S2† )c + H.c.. (7) Ha,c = 2


Dissipation, which can occur via the fiber loss, atomic spontaneous emission and cavity decay, is a requisite component in the current scheme. The states in two-excitation subspace would be transformed to the corresponding states in one-excitation and zero excitation subspace via dissipation. The dynamics of the open dissipative system in Lindblad form could be described by the master equation 1 ˆ ˆ† (8) 2 L j ρˆ L j − ( Lˆ †j Lˆ j ρˆ + ρˆ Lˆ †j Lˆ j ) , ρ˙ˆ = i[ ρ, ˆ H] + 2 j Specifically, in the current where Lˆ j is the so-called Lindblad operators governing dissipation. √ √ √ scheme the Lindblad operators can be expressed as Lˆ β = βb, Lˆ κ1 = κa A , Lˆ κ2 = κa B , √ √ ˆ ˆ Lˆ γ1 = γ|g √ A A e| and Lγ2 = √ γ|g BB e|. L β describes the dissipation induced by the fiber loss. Lˆ κ1 = κa A and Lˆ κ2 = κa B describe the dissipation induced by the leakage of cavity A √ √ and cavity B, respectively. Lˆ γ1 = γ|g A A e| and Lˆ γ2 = γ|g BB e| describe the dissipation induced by the spontaneous emission of atom in cavity A and B, respectively. Since cavity A and cavity B are distant from each other, the dissipation processes are spatially separated. 3. 3.1.

Preparation of the distributed entanglement Dressed states Ψ 22,+ Ψ +2

Φ

Ψ2,2 −

Φ +2

0 2

3g Φ21

3g Φ

ω0

ω2

− 2

ω2

Ψ 12,+ Ψ−2 Ψ 2,1 −

Φ +1

Φ 01

Φ 22

Ψ1+

2g Φ −1

ω0

Ψ1−

ω1 Φ0

Fig. 2. Schematic diagram for the dressed state of the atom-cavity-fiber coupling system.

The spectroscopy of the system is well described by the dressed states, i.e., the eigenstates of the Hamiltonian H0 + Hc, f + Ha,c , as shown in Fig. 2. Here, we use them to see clearly the roles of Hcl after choosing appropriate laser driving and detuning. We define the excitation number (|eii e| + ai† ai ) + b† b. In Table 1, the eigenstates and operator of the total system Ne = i=1,2

corresponding eigenvalues of zero and single excitation subspaces are shown with the notation |Φ0 = |g1 , g2 |0, 0, 0,

|Φ01 = |φ+ |0, 0, 0,


|Ψ1− = |g1 , g2 |0, 0, 1, |Ψ1+ = |g1 , g2 |0, 1, 0, 1 |Φ1± = √ [|φ − |0, 0, 0 ± |g1 , g2 |1, 0, 0], 2

(9)

√ with |φ± = 1/ 2(|e1 , g2 ± |g1 , e2 ). In Table 2, the eigenstates and corresponding eigenvalues Table 1. The eigenstates and eigenenergies of the Hamiltonian H0 + Hc, f + Ha,c within the zero and single excitation subspaces.

Eigenstate |Φ0 |Φ01 |Ψ1± |Φ1±

Eigenenergy 0 ω0√ ω0 ± 2ν ω0 ± g

of two excitation subspace are shown with the notation |Ψ2− = |φ+ |0, 0, 1, |Ψ2+ = |φ+ |0, 1, 0, 1 1 |Ψ2,± = √ [|φ − |0, 0, 1 ± |g1 , g2 |1, 0, 1], 2 1 2 = √ [|φ − |0, 1, 0 ± |g1 , g2 |1, 1, 0], |Ψ2,± 2 |Φ02 = |φ+ |1, 0, 0, |Φ12 = |g1 , g2 |0, 1, 1, √ 1 |Φ22 = √ (|g1 , g2 |2, 0, 0 + 2|e1 , e2 )|0, 0, 0, 3 1 1 √ |Φ2± = √ [(|φ − )|1, 0, 0 ± √ ( 2|g1 , g2 |2, 0, 0 3 2 −|e1 , e2 |0, 0, 0)].

(10)

Table 2. The eigenstates and eigenenergies of the Hamiltonian H0 + Hc, f + Ha,c within the two-excitation subspace.

Eigenstate |Ψ2± 1 |Ψ2,± 2 |Ψ2,± |Φ02 |Φ12 |Φ22 |Φ2±

Eigenenergy √ 2ω0 ± 2ν √ 2ω0 ± g − 2ν √ 2ω0 ± g + 2ν 2ω0 2ω0 2ω0√ 2ω0 ± 3g

Under the weak excitation condition and if the initial state is in the zero excitation subspace, we can safely discard the subspace with excitation number greater than or equals two, and |Φ01 is the maximal entanglement we want to prepare for the two distributed atoms.


3.2.

Roles of the classical laser field

Under the dressed state picture, The Hamiltonian Hcl can be rewritten as 1 1 Ωk e −iω k t (|Φ1− + |Φ+1 )Φ0 | + √ Ωk e −iω k t |Φ01 Φ0 | Hcl = 2 2 k=1,2 1 1 + √ Ωk e −iω k t |Φ2− Φ1− | + Ωk e −iω k t (|Φ12,− + |Φ12,+ )Φ1− | 2 2 1 1 + √ Ωk e −iω k t |Φ+2 Φ+1 | + Ωk e −iω k t (|Φ22,− + |Φ22,+ )Φ+1 | 2 2 1 1 − √ Ωk e −iω k t |Φ22 + Ωk e −iω k t |Φ02 (Φ1− | + Φ+1 |) 2 6 √ 2 + Ωk e −iω k t (|Φ2− + |Φ+2 )(Φ1− | − Φ+1 |) 4 1 −iω k t 2 0 |Φ2 Φ1 | + H.c. . + √ Ωk e 3

(11)

We focus the discussion on the interaction picture with respect to the Hamiltonian H0 + Hc, f + Ha,c that expressed by the eigenvectors and eigenvalues in zero, one and two excitation subspace. By choosing the detunings Δ1 = ω0 − ω1 and Δ2 = ω2 − ω0 equal to g, only five resonant transitions can occur, i.e. |Φ0 → |Φ1− , |Φ+1 → |Φ02 and |Φ+1 → |Φ22 induced by the driving field Ω1 , |Φ1− → |Φ02 and |Φ1− → |Φ22 driven by the driving field Ω2 , while all other transitions between arbitrary two dressed states are largely detuned. 3.3.

Preparation process Φ 22

Φ 02

κ

ω1

ω2

ω1

κ

ω2

κ Φ1+

Target state

Φ10

Φ1−

κ,γ γ

ω1

κ,γ Φ0

Fig. 3. Processes for producing and stabilizing Bell state. The interaction between system and environment is characterized by the photon loss and atomic spontaneous emission with the rates κ and γ, respectively.

The processes for producing and stabilizing the Bell state are shown in Fig. 3, we here require the atomic spontaneous emission to be much slower than other dynamical processes. Besides, the generation of the steady-state is independent of the initial states. If the system is initially in


the ground state |Φ0 , it will be driven by the classical field Ω1 to one-excitation dressed state |Φ1− and then to |Φ02 and |Φ22 by classical field Ω2 . The photon loss results in the decaying channel |Φ02 → |Φ01 = |φ+ |0, 0, 0. The state |Φ01 is decoupled from the qubit-resonator coupling and cavity decay and is unaffected by the drives due to off-resonant, so it is a steady-state. On the other hand, the state |Φ22 decays to the one-excitation dressed state |Φ1± , repumped by the classical field Ω1 to |Φ02 and |Φ22 . With the coherent driving and dissipation processes continuing, the population of dressed state |Φ22 will decline gradually until all of qubit population is driven to the Bell state |Φ01 . If the system is initially in |e1 , g2 |0, 0, 0 or |g1 , e2 |0, 0, 0, which can be regarded as a superposition of the one-excitation dressed states |Φ01 and |Φ1± . As has been shown, the populations of |Φ1± are finally transferred to the steady-state due to coherent driving and dissipation process. For initial state |e1 , e2 |0, 0, 0, the strong qubit-resonator coupling results in the transfer of population to the states |φ− |1, 0, 0 and |g1 , g2 |2, 0, 0. Due to dissipation, these two state continuously decay to |φ− |0, 0, 0 and |g1 , g2 |1, 0, 0, respectively, the specific superposition of the dressed states |Φ+1 and |Φ1− , which will be driven to the steady-state finally. 3.4.

Performance of the scheme

In this part, we present numerical simulation of the full dynamics of the whole system. The most common way to assess the quality of the steady-state is fidelity which is defined as F (t) =Tr[(|φ+ φ+ | ⊗ Ic ) ρˆ t →∞ ] with |φ+ being the desired state and ρˆ t →∞ being the practical steady-state density matrix. The premise of locality and realism implies some constraints on the statistics of two spatially separated particles, which is known as Bell inequalities [50]. On that basis, CHSH correlation [51]. The system state violates Bell inequality when the CHSH √ correlation rises above 2, and quantum mechanisms predicts CHSH correlation S(t) equals 2 2 for the maximal violation limit. The CHSH correlation S(t) is defined as S(t) = Tr[(OCHSH ) ρ(t)],

(12)

with respect to evolution time, where OCHSH is defined as OCHSH = σ y,1 ⊗ +σ x,1

−σ y,2 − σ x,2 −σ y,2 − σ x,2 + σ x,1 ⊗ √ √ 2 2 σy,2 − σ x,2 σy,2 − σ x,2 ⊗ − σ y,1 ⊗ . √ √ 2 2

(13)

We also introduce purity to characterize mixture degree of the target steady-state entanglement which is defined through reduced density operators as P (t) = Tr[ ρ(t) ˆ 2 ].

(14)

If the system is in a pure state, the purity is precisely unit. To verify the feasibility of the scheme, we set |Φ0 = |g1 , g1 |0, 0, 0 as the initial state, and solve the master equation numerically in zero, one and two excitation subspace. In our scheme, the decay rate of cavity field modes need to be far larger than the spontaneous emission rate of qubits, i.e. γ κ, the coupling to the cavity modes should be much stronger than the coupling to the driving field, i.e. Ω1 , Ω2 g, and the decay rate of cavity field modes and fiber mode should be comparable with weak driving strengths, i.e. κ, β Ω1 , Ω2 . With these assumptions, we can efficiently generate and protect the Bell state |φ+ . We choose a set of feasible experimental parameters in a recent circuit QED experiment [52]: χ/2π 6 MHz, κ/2π 1.7 MHz, T1 9 μs, where T1 is the qubit energy relaxation time, and χ = g2 /Δ, with Δ being the qubit-resonator

1

1

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X: 6000 Y: 0.9153

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Fidelity


X: 6000 Y: 0.9724

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0.6

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gt

gt

(a)

(b)

4000

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Fig. 4. The fidelity of states |Φ0 and |Φ01 versus the dimensionless parameter gt for the initial state |Φ0 by solving the full master equation. In (a) the experimental parameters are chosen as Ω1 = 0.080g, Ω2 = 0.035g, ν = 20g. And the dissipative factors are chosen as γ/g = 2.72 × 10 −4 , κ 1 = κ 2 = 2.8 × 10 −2 g, β = 1.5 × 10 −2 g. The inset of (b) is plotted with the optimized parameters: Ω1 = 0.025g, Ω2 = 0.045g and ν = 20g. And the dissipative factors: γ = 1.75 × 10 −5 g, κ 1 = 1.6 × 10 −2 g, κ 2 = 2.4 × 10 −2 g and β = 0.67 × 10 −2 g.

detuning. It is reasonable to set Δ = 10g in this dispersive qubit-resonator interaction system, yielding g/2π 60M H z, κ 1 = κ 2 2.8 × 10 −2 g, β = 1.5 × 10 −2 g and γ 2.72 × 10 −4 g. The optimized Rabi frequencies are taken as Ω1 = 0.080g and Ω2 = 0.035g. These parameters are feasible in experiment for a number of quantum optical and solid state systems, for instance: ion traps systems [44, 49], superconducting circuit QED systems [52], and plasmonic systems [53– 55]. Especially, for the future implementation of quantum information with cold atoms, since the qubits need to be confined within the waist of the cavity mode in cavity QED model, and this can be realized with cold atoms trapped in a far detuned optical dipole trap [56, 57] or optical lattice [58–60], which can be mapped into a cavity QED model as our scheme utilized. With these parameters, in Fig. 4(a) we plot the evolutions of the experimental fidelity versus the steady-state |Φ01 and initial state |Φ0 . The result shows that the desired state could be prepared with fidelity more than 90% when the evolution time equals 1500/g. In Fig. 4(b), we optimized the dissipative factors and Rabi frequencies of the drivings by taking γ = 1.75 × 10 −5 g, κ 1 = 1.6 × 10 −2 g, κ 2 = 2.8 × 10 −2 g, β = 0.67 × 10 −2 g, Ω1 = 0.025g, Ω2 = 0.045g, with these optimized parameters, the optimal fidelity for steady Bell state is about 97.24% when the evolution time equals 3500/g. It is shown that the evolution time needed to arrive at the steadystate increases with the decreasing of Ω1 and Ω2 and the corresponding fidelity decreases with the increasing of γ. This can be explained as follows, when Ω1 and Ω2 are decreased, the unitary dynamics becomes slower and the time to reach stabilized state increases. On the other hand, the transition from the Bell state to the ground state is strongly suppressed as γ decreases, thus the fidelity increases. With the experimental parameters and optimized parameters mentioned above, we plot the CHSH correlation and the purity of the target steady-state as a function of the evolution time, as shown in Fig. 5. The Numerical simulation in Fig. 5(a) shows that the system is reasonably stabilized to the target steady-state, with the experimental CHSH correlation about 2.532 and the optimal CHSH correlation about 2.731, which are both clearly exceeding the maximum value of 2 allowed by the local hidden variable theories. One can see from Fig. 5(b) that the evolution curve of experimental purity and optimal purity exhibit a valleys in the regime 0< t