Abstract
Quantum vacuum forces dictate the interaction between individual atoms and dielectric surfaces at nanoscale distances. For example, their large strengths typically overwhelm externally applied forces, which makes it challenging to controllably interface cold atoms with nearby nanophotonic systems. Here we theoretically show that it is possible to tailor the vacuum forces themselves to provide strong trapping potentials. Our proposed trapping scheme takes advantage of the attractive groundstate potential and adiabatic dressing with an excited state whose potential is engineered to be resonantly enhanced and repulsive. This procedure yields a strong metastable trap, with the fraction of excitedstate population scaling inversely with the quality factor of the resonance of the dielectric structure. We analyse realistic limitations to the trap lifetime and discuss possible applications that might emerge from the large trap depths and nanoscale confinement.
Introduction
One of the spectacular predictions of quantum electrodynamics is the emergence of forces that arise purely from quantum fluctuations of the electromagnetic vacuum^{1}. Also known as Londonvan der Waals^{2} or Casimir forces^{3} in different regimes, these forces are often dominant at short distances and can give rise to undesirable effects in nanoscale systems, such as nanomechanical stiction^{4}. These forces have also attracted increasing attention in the fields of atomic physics and quantum optics. In particular, significant efforts have been made in recent years to interface cold atoms with the evanescent fields of dielectric micro and nanophotonic systems^{5,6,7,8,9,10,11,12,13,14}. These systems are expected to facilitate strong, tunable interactions between individual atoms and photons for applications such as quantum information processing^{15} and the investigation of quantum manybody physics^{16,17,18}. In practice, efficient atomic coupling to the evanescent fields of the nanophotonic systems requires that atoms be trapped within subwavelength distances of these structures.
At these scales, quantum vacuum forces can overwhelm the forces associated with conventional optical dipole traps, typically resulting in a loss of trap stability at distances d≲100 nm from dielectric surfaces. Given the ability to engineer the properties of nanophotonic structures, an interesting question arises as to whether such systems could be used to significantly modify vacuum forces, perhaps changing their sign from being attractive to repulsive, or even creating local potential minima. The strength of nanoscale vacuum forces should lead to unprecedented energy and length scales for atomic traps, which would find use beyond nanophotonic interfaces, such as in quantum simulation protocols based on ultracold atoms^{19} and control of interatomic interactions^{20}.
Here we propose a novel mechanism in which engineered vacuum forces can enable the formation of a nanoscale atomic trap. While a recent nogo theorem^{21} forbids a vacuum trap for atoms in their electronic ground states, tailored nanophotonic systems can yield strong repulsive potentials for atomic excited states. We show that a weak external optical field can give rise to an overall trapping potential for a dressed state. Remarkably, absent fundamental limits on the losses of the surrounding dielectric structure, the fraction of excitedstate population in the dressed state can become infinitesimal, which greatly enhances the trapping lifetime and stability. We identify and carefully analyse the actual limiting mechanisms. While we present calculations on a simple model where analytical results can be obtained, we also discuss the generality of our protocol to realistic systems such as photonic crystal structures.
Results
Quantum interactions between atom and surface
The nogo theorem for nonmagnetic media states that a dielectric object in vacuum cannot be stably trapped with vacuum forces for any surrounding configuration of dielectric objects, provided that the system is in thermal equilibrium^{21}. In analogy with Earnshaw’s theorem, which prohibits trapping of charged objects with static electric potentials, at best one can create a saddlepoint potential (see, for example, ref. 22 involving metal particles and refs 13, 23 for a hybrid vacuum/optical trap for atoms). For dielectric objects, potential loopholes involve embedding the system in a highindex fluid^{24} or using blackbody radiation pressure associated with strong temperature gradients^{25}. Applied to atoms, this theorem essentially forbids the stable vacuum trapping near a dielectric structure when the atom is in its electronic ground state. No such constraint exists for atoms in their excited states^{26}, although the robustness of any possible trap would be limited by the excitedstate lifetime. Here, we show that an atom only weakly dressed by its excited state can be trapped by properly tailoring the dispersion of the underlying structure.
We begin by briefly reviewing how quantum vacuum fluctuations give rise to forces on the ground and excited states. Within an effective twolevel approximation of an isotropic atom with ground and excited states g›, e›, respectively, the interaction between the electromagnetic field and an atom at position r is given by the dipole Hamiltonian H=−d·E(r). Conceptually, following a complete decomposition of the electric field into its normal modes k, one can write , where . The energy nonconserving terms ( and ) enable an atom in its ground state g, 0› to couple virtually to the excited state and create a photon, e, 1_{k}›, which is subsequently reabsorbed. The corresponding frequency shift for the ground state within secondorder perturbation theory is δω_{g}(r)=−∑_{k} g_{k}(r)^{2}/(ω_{0}+ω_{k}), where ω_{0} is the unperturbed atomic transition frequency. This shift results in a vacuuminduced mechanical potential when translational symmetry is broken owing to dielectric surfaces. Using a quantization technique for electrodynamics in the presence of dispersive dielectric media, the shift can be expressed in terms of the scattered component of the dyadic electromagnetic Green’s function evaluated at imaginary frequencies ω=iu (ref. 27) (also see Methods),
where Γ_{0} is the freespace spontaneous emission rate of the atom. For simplicity, we have given the result in the zerotemperature limit. Finitetemperature corrections^{28,29} associated with atomic interactions with thermal photons are only relevant for distances from surfaces comparable with or larger than the thermal blackbody wavelength, d≳λ_{T}, where at T=300 K.
A similar process can occur for an atom in its excited state e, 0›, in which the atom virtually emits and reabsorbs an offresonant photon (thus coupling to the state g, 1_{k}›). However, the excited state is unique in that it can also emit a resonant photon. The total excitedstate shift is given by^{27}
The first term arises from offresonant processes, while the second term can be interpreted as the interaction between the atom and its own resonantly emitted photon. Naturally, the dielectric environment can modify the spontaneous emission rate as well,
Note that the resonant shifts and modified emission rates are complementary, in that they emerge from the real and imaginary parts of the Green’s function. Micro and nanophotonic systems are already widely exploited to modify emission rates of atoms and molecules in a variety of contexts^{30,31,32,33}. We propose that these same systems can be used to tailor excitedstate potentials to facilitate vacuum trapping.
Interactions with a Drude material
In what follows, we present a simple model system that enables onedimensional trapping in a plane parallel to a semiinfinite dielectric slab, as illustrated in Fig. 1. The use of a homogeneous dielectric (as opposed to, for example, a photonic crystal) greatly simplifies the calculation and enables one to understand the relevant trap properties analytically, although we later argue why the qualitative features should still hold in more complex, realistic settings. In the dielectric slab model, dispersion engineering will be realized via the frequencydependent electric permittivity ε(ω). While this particular system achieves a onedimensional trap along z, our electrodynamic calculations are performed in three dimensions.
A shortdistance expansion of the Green’s function, valid for subwavelength scales, shows that the excitedstate emission rate Γ(z) and resonant contribution to the shift are given by
where the permittivity ε_{a}=ε(ω_{0}) is evaluated at the atomic resonance frequency, z is the distance from the surface and k_{0}=ω_{0}/c is the resonant freespace wavevector. A large repulsive potential is generated when ε_{a} approaches −1 from above, ε_{a}→−1^{+}, which is analogous to the interaction between a classical dipole at position z and its large induced image dipole in the dielectric. The modified spontaneous emission rate originates from the absorption or quenching of the atomic emission owing to material losses.
We choose a Drude model, , as a simple dielectric function. Note that this function satisfies Kramers–Kronig relations (causality), as a choice of a noncausal function could result in an apparent violation of the nogo theorem^{21} as well. In the limit of vanishing material loss parameter γ, the system passes through ε=−1 at the plasmon resonance frequency . The system is conveniently parameterized by the quality factor Q≡ω_{pl}/γ and a dimensionless detuning Δ_{p}=(ω−ω_{pl})/γ. For Q>>Δ_{p}, the factor resembles the complex susceptibility of a simple resonator, which in the fardetuned limit (Δ_{p}>>1) yields and . The dispersion and dissipation scale like and , respectively. Thus, for high Q, it is possible to choose detunings where the atom still sees significant repulsive forces, but where spontaneous emission into the material is not significantly enhanced (in addition to materialinduced emission of equation (4), there still is emission into freespace at a rate ∼Γ_{0}). This scaling behaviour is in close analogy with conventional optical trapping of atoms, where simultaneously using large trapping intensity and detuning maintains reasonable trap depths but suppresses unwanted photon scattering. The groundstate shift for atomic frequencies ω_{0}∼ω_{pl} and in the nearfield is given by . An additional condition for the validity of the shortdistance results obtained for the excited state is , as shown in the Methods. We continue to present approximate results in this regime owing to its simplicity, although all of our numerical results (such as in Figs 2c,d and 3) are calculated using the full Green’s functions.
We also briefly comment on two effects in realistic systems that can be safely excluded from our analysis. First, in principle, the excitedstate shift can be affected by coupling to even higher electronic levels. However, a careful calculation of this contribution^{34} shows that it is negligible compared with the resonantly enhanced term of equation (5), provided that we work in the regime Q>>Δ_{p}. Second, we have ignored corrections to the shifts owing to surface roughness. It can be shown that the length scale for modifications to the shifts is roughly set by the size of the surface variation itself^{35}. The realistic regimes of operation of our trap will be at d≳10 nm from surfaces, which allows us to safely ignore the angstromscale surface roughness achievable in stateoftheart dielectric structures^{36}.
Formation of a vacuum trap
We now describe how a trap can form for an atom subject to these potentials, and in the presence of a weak driving laser of frequency ω_{L} and Rabi frequency Ω (see Fig. 1). For conceptual simplicity, we choose the Rabi frequency to be spatially uniform, such that all atomic forces are attributable to the vacuum potentials alone. Qualitatively, the strong position dependence of the ground and excitedstate frequencies causes the laser to come into resonance with the atom at a single, tunable point z=z_{b}. Under certain conditions (described in detail later), an atom starting at position z>>z_{b} will be far detuned, such that it is essentially in the ground state and is attracted by the pure groundstate potential. However, moving closer to z_{b} brings the atom closer to resonance. The dressing of the atom with a small fraction of excitedstate population causes a repulsive barrier to form near z∼z_{b}, yielding a metastable trap. Significantly, our approach does not attempt to directly counteract the attractive groundstate potential with large external optical potentials, in sharp contrast with other trapping schemes near dielectric surfaces^{8,9,11,12,37,38,39}.
The order of our calculation of the trap properties is as follows. We first find the adiabatic potential experienced by a slowly moving atom. We derive relevant properties around the location of the trap minimum z_{t} (Fig. 1). In particular, we show that absent material constraints (arbitrarily high Q), the dressed state can in principle have infinitesimal excitedstate population and scattering. We then quantize the motion to find the motional eigenstates, binding energy and position uncertainty Δz. Finally, we analyse the mechanisms that limit the trap lifetime.
Following standard procedures^{40}, the average force experienced by the atom is given by the expectation value of the Heisenberg equation , while the internal atomic dynamics satisfies the usual Bloch equations, for example,
Here, for notational simplicity, denotes the expectation value of the atomic variables, and we have defined δ_{a}(z)=ω_{L}−(ω_{e}(z)−ω_{g}(z)) as the detuning between the laser and the local atomic resonance frequency. When the atom moves slowly on the timescales of the internal dynamics, the atomic coherence can be adiabatically eliminated, , such that the atomic populations are functions of position alone, . We are primarily interested in the regime where the atom is weakly driven, σ_{ee}<<1. A more careful calculation reveals that the Green’s function G_{sc}(r, r, ω) for the excitedstate resonant shift and emission rate in equations (2) and (3) must in fact be evaluated at the laser frequency in this regime, which reflects that the atom primarily acts as a Rayleigh scatterer (see Methods). Consequently, ε_{a} in equations (4) and (5) is replaced with ε(ω_{L}). The adiabatic potential seen by a slowly moving atom is . In principle, the Heisenberg equation for should also include the effects of spontaneous emission. It can be shown, however, that this term contributes a zero average force (see Methods), and only results in a heating term to be analysed later.
Three independent frequencies characterize the system (ω_{L}, ω_{p} and ω_{0}), and two relative frequencies must be specified. One is determined by selecting the trap barrier position, determined by the relation δ_{a}(z_{b})≡0. A second is the dimensionless detuning between the plasmon resonance and laser, Δ_{p}=(ω_{L}−ω_{pl})/γ, which we treat as a free parameter that determines the relative strengths of the excitedstate shift and dissipation. In the nearfield, the excitedstate repulsion exceeds the groundstate attraction by a factor 2Q/3Δ_{p}. The population at z_{b} must then exceed the inverse of this quantity, σ_{ee}>3Δ_{p}/2Q, to provide a classical barrier in the adiabatic potential, which translates to a minimum Rabi frequency . Since Δ_{p}≲Q, this scheme potentially enables atom trapping near surfaces with greatly reduced intensities as compared with conventional trapping. In the latter case, an optical dipole potential requires a minimum driving field of to overcome the attractive groundstate potential.
The excitedstate population at the trap minimum is fixed by the ratio of ground to excitedstate forces, , and is surprisingly independent of Ω (provided that Ω>Ω_{min}). An increasing Rabi frequency has the effect of increasing the local detuning δ_{a}(z_{t}) to preserve the same population. This manifests itself as a selfselection of the trapping position and an increasing distance from the barrier with increasing intensity,
where we have defined the dimensionless Rabi frequency . The ability to tune the trap position using both the laser amplitude and its frequency (which determines z_{b}) enables the adiabatic transformation between deep but small traps close to the surface to larger, shallow traps ∼100 nm from the surface. This is expected to greatly facilitate trap loading, as techniques to trap atoms at these larger distances are already well established^{8,11,12,41}.
The photonscattering rate at the trap position, R_{sc}=Γ(z_{t})σ_{ee}(z_{t}), leads to lifetime limitations of our trapping scheme. Fixing other parameters, the detuning Δ_{p} from the plasmon resonance can be optimized to yield the minimum scattering rate, . Thus, absent fundamental limits on the magnitude of Q, the photonscattering rate can in principle be suppressed to an arbitrary degree. This implies that an infinitesimal violation of the assumptions underlying the nogo theorem can in fact allow for trapping.
In a conventional optical dipole trap, the potential U=−α(ω_{L})E(r, ω_{L})^{2}/2 seen by an atom depends on the local field intensity and is proportional to a spatially constant atomic polarizability evaluated at the laser frequency. In contrast, here we exploit a novel mechanism arising from strong spatial variations in the polarizability itself, induced through the vacuum level shifts. This creates a backaction effect analogous to that producing a dynamical ‘optical spring’ in optomechanical systems^{42}, and enables extremely steep trap barriers around z_{b} (as in Fig. 1c). In particular, the repulsive force is given by , where is engineered to be large. However, the excitedstate population itself varies strongly with position through the spatially dependent detuning (for simplicity, here we omit the varying linewidth Γ(z)). Taking into account the spatial variation of σ_{ee} around the equilibrium position, one finds an overall repulsive force that scales like ,
Results for trapping of caesium
Given the possible steepness of this barrier, an ideal limit for the overall potential is shown in Fig. 2a. The attractive part consists of the pure groundstate potential and arises from quantum fluctuations, while an infinite hard wall is created at z=z_{b} owing to the combination of strong resonant excitedstate shifts (originating from the interaction of the atom with a classical image dipole) and optomechanical backaction. This idealized model enables one to understand the best possible scaling of the trap properties. In particular, the trap depth U_{depth} (the energy required for a classical particle to become unbound) cannot exceed the value of the groundstate potential at z_{b}, U_{max}≡ħδω_{g}(z_{b}) (for example, U_{max}≈60 mK when z_{b}=10 nm for a Caesium atom). This trapping potential can be easily quantized numerically. In Fig. 2b, we plot the quantumbinding energy U_{0} (where 0<U_{0}<U_{depth}≤U_{max}) and position uncertainty Δz for the motional groundstate wavefunction. For the numerical results in Figs 2 and 3, we use atomic properties corresponding to Caesium (λ_{0}=852 nm, Γ_{0}/2π=5.2 MHz, recoil frequency ω_{r}/Γ_{0}=4 × 10^{−4}). The strength of the vacuum potentials is reflected in the strong confinement of the wavefunctions (for example, Δz<4 nm for distances z_{b}<50 nm), which are an order of magnitude smaller than in conventional optical traps. To contrast our results for Caesium with those of a less massive atom, we have considered Lithium as well. Results for Li similar to those for Cs in Fig. 2a,b show that at z_{b}=10 nm, the potential depth is smaller by roughly a factor of two, while the confinement Δz sees roughly a fivefold increase. Generally, within the Wentzel–Kramers–Brillouin approximation and to lowest order in Δz, the groundstate localization scales like k_{0}Δz∼(ω_{r}(k_{0}z_{b})^{4}/Γ_{0})^{1/3}, while the binding energy U_{0}≈U_{max}(1−3Δz/z_{b}) can be nearly the entire classical potential depth.
This ideal model describes qualitatively well the actual potentials over a large parameter regime. In Fig. 2c, we plot the ratio of the numerically obtained trap depth to the theoretical maximum, U_{depth}/U_{max}, as functions of Δ_{p} and dimensionless driving amplitude . Here, we have chosen a quality factor of Q=10^{7} and a barrier position of z_{b}=10 nm. A trap depth comparable with U_{max} is achievable over a large range of values. For low values of , the trap depth is weakened owing to escape over the potential barrier, while for large values, the trap becomes shallow as its position z_{t} is pulled away from z_{b} (as discussed previously). Two representative trap potentials, along with the groundstate wavefunctions and confinement Δz are illustrated in Fig. 2d.
Analysis of trap heating
We have identified four sources of motional heating—recoil heating, nonadiabatic motion of the atom within the trap, fluctuations of the atomic coherence and tunnelling over the finiteheight potential barrier—that limit the trap lifetime in absence of external cooling mechanisms. Tunnelling is exponentially suppressed with barrier height and thus is easily suppressed with increasing Rabi frequency. This calculation can be found in the Methods, while we discuss the more important mechanisms here.
As in conventional optical trapping, the random momentum recoil imparted on the atom by scattered photons yields an increase in motional energy at the rate . Two qualitative differences emerge relative to free space, however. First, the photonscattering rate is modified in the presence of dielectric surfaces. Second, the effective momentum ħk_{eff} imparted by a photon is enhanced compared with the freespace momentum. This momentum scales at close distances like (see Methods), owing to emission of highwavevector photons into the dielectric. Starting from the motional ground state, this heating causes the atom to become unbound over a timescale τ_{r}≈(k_{0}z_{t})^{2}Δ_{p}/(3ω_{r}). It should be noted that the heating rate dE/dt does not depend on the atom confinement (that is, whether we operate in the LambDicke limit), as the suppression of motional jumps in tight traps is compensated by the increase in energy gained per jump^{43}. The increase in lifetime with trap position, , is attributable entirely to the reduction in recoil energy associated with a photon emitted into the surface. In particular, the decreasing trap depth versus distance is exactly compensated for by the reduced surfaceenhanced emission rate .
Our derivation of the atomic potential assumes that the atomic coherence adjusts rapidly to the local trap properties, dσ_{ge}/dt≈0, which results in a purely conservative potential. As we now show, nonadiabatic corrections result in a momentumdependent force, dp/dt∝p (antidamping). We can solve equation (6) for the atomic motion perturbatively by writing , where is the adiabatic solution. It depends only on the instantaneous position and is given by in the weak saturation limit. Here, δ_{c}(z)=δ_{a}(z)+iΓ(z)/2 is the complex detuning. Substituting this solution into equation (6) yields a velocitydependent correction to the coherence, . This in turn yields a new contribution to the force, F^{(1)}=βp, where the antidamping rate is given by evaluated at the trap position z_{t}. Intuitively, the atom is antidamped (β>0) because the laser frequency is bluedetuned relative to the atomic frequency at z_{t}, giving rise to preferential Stokes scattering over antiStokes. Aside from the term proportional to dΓ/dz, the heating rate is exactly analogous to that occurring in an optomechanical system in the bluedetuned regime^{42}. We define a characteristic time τ_{ad} needed for the energy increase produced by antidamping to equal the groundstatebinding energy U_{0}.
Thus far, we have focused on the motion of the atomicdressed state, in which a trap emerges owing to weak mixing with the excitedstate potential. Following a spontaneous emission event, however, the atom returns to the ground state and sees the pure groundstate potential for a transient time t_{trans}∼δ_{c}(z_{t})^{−1} until the atomic coherence is restored, during which the atomic wave packet accelerates toward the surface and gains energy. While a simulation of the full dynamics of the atom (including the spatial and internal degrees of freedom) is quite challenging, here we estimate the heating rate from these transient processes. In particular, we first find the spatial propagator under the pure groundstate potential, U(z_{f}, z_{i}, t_{f}, t_{i}) (see Methods). Then, assuming that a spontaneous emission event occurs at t_{i}=0 with the atom initially in the motional ground state ψ(z_{i}), the new wavefunction that emerges once the coherence equilibrates is given by ψ(z_{f})∼∫dz_{i}U(z_{f}, z_{i}, t_{trans}, 0)ψ(z_{i}). The probability of atom loss P_{trans} per spontaneous emission event is estimated from the overlap between the initial and final states, P_{trans}=1−‹z_{f}z_{i}›^{2}, with a corresponding trap lifetime of .
The overall trap lifetime is obtained by adding the individual lifetimes in parallel, , and is plotted in Fig. 2a. The overall trap lifetime exhibits a complicated dependence on Rabi frequency Ω and detuning Δ_{p} owing to the different scalings of the constituent heating mechanisms (see Fig. 3b). As an example, however, for external parameters Δ_{p}=2.5 × 10^{5} and Ω/Ω_{min}=1.5 × 10^{4} for atomic Cs, one achieves a lifetime of τ_{total}≈15 ms, groundstate confinement of Δz≈1 nm, trap depth of U_{depth}≈10 mK and trap distance of z_{t}≈15 nm. This Rabi frequency corresponds to an incident intensity of just 0.05 mW μm^{−2}. To compare, nanofibrebased optical traps^{8,11} employ guided intensities three orders of magnitude larger to achieve confinement and surface distances of Δz∼20 nm (corresponding to trap frequencies of ) and z_{t}∼200 nm, respectively. Trapping frequencies of ω_{m}/2π≈40 MHz would be needed to reach singlenanometer confinement, far beyond what is realistic in optical traps.
An additional heating mechanism, which does not appear in our simple model of an isotropic twolevel atom, arises from the possibility of an excited state in a realistic multilevel atom to emit photons of different polarizations and transition into different states in a groundstate manifold. The different Clebsch–Gordan coefficients of these transitions, along with the symmetry breaking introduced by the nearby dielectric structure, imply that different excited states can experience different vacuum shifts. This would appear as different trapping potentials for the various dressed groundstate sublevels absent of some special engineering. As a result, spontaneous changes in the sublevels (such as arising from spontaneous photon scattering) would induce fluctuations of the trapping potential and lead to motional heating. One possibility to prevent these issues in an experiment is to utilize the effective twolevel structure provided by a cycling transition.
Discussion
We have described a protocol for an atomic trap based on engineered vacuum forces, and have analysed in detail a model case of onedimensional trapping near a Drude material, where analytical results are possible to obtain. The Drude response approximates well a number of metals, such as silver and gold, in the optical domain^{44}. In practice, however, these metals are limited by their lowquality factors (Q<10^{2}), and the thermal fluctuations of such conducting materials can give rise to strong noiseinduced trap heating and decoherence^{45}. On the other hand, photonic crystal structures enable engineering of resonances through geometry, and quality factors approaching Q∼10^{7} have been observed^{46}. We anticipate that our trapping protocol could be quite generally applied, and the resulting trap behaviour would qualitatively remain the same. In particular, the spatial dependence of the excitedstate properties would no longer scale like (k_{0}z)^{−3} as in the case of a simple dielectric surface, but be characterized by a more complex spatial function f(r) that reflects the mode shape of the underlying structure^{23}. However, the scaling of the excitedstate shift (δω_{e}∝Q/Δ_{p}) and emission () is generic to coupling between an atom and resonator of Lorentzian lineshape, and thus the capability of achieving highly stable traps with highQ structures using our proposed techniques is maintained. Photonic crystals thus offer great flexibility in tailoring the properties of vacuum traps (such as dimensionality), which we plan to investigate in detail in future work. For example, it has been shown that dielectric gratings can produce quantum vacuum forces that trap atoms along the lateral directions of a dielectric interface^{35}, which when combined with our previous techniques could yield full threedimensional traps. We note that theoretical^{23} and experimental efforts^{13} have already been made to achieve realistic hybrid atom traps in photonic crystal structures, where a combination of vacuum forces and optical forces is required to stabilize the trap in all dimensions. In addition, while we have focused on trapping of atoms weakly dressed by an external laser, it would be interesting to investigate the feasibility of laserfree traps, where atoms pumped into metastable electronic levels are stably confined using engineered resonant shifts alone.
The ability to create atomic traps with parameter sets (such as depth, confinement and proximity to surfaces) that are not possible in conventional traps (whether optical, electrical or magnetic) should lead to a number of intriguing applications. For example, it has been proposed that the trapping of atoms near dielectric surfaces with nanoscale features (such as a subwavelength lattice) would facilitate large interaction strengths and longrange interactions for ultracold atoms^{39}. However, the major limitation of optical trapping is the divergent intensity needed to overcome the groundstate vacuum forces of these structures. The use of resonantly enhanced excitedstate repulsion to create a trap could significantly reduce the power requirements and enable smaller lattice constants. Furthermore, traps with localization on the nanometer scale can induce interatomic forces that are of comparable strength with molecular van der Waals forces, which enables novel opportunities to control ultracold atomic collisions^{20} and realize exotic interactions such as pwave scattering of fermions^{47}. Atoms trapped near surfaces could also act as exquisite nanoscale probes of surface physics, including the precision measurement of vacuum forces themselves.
Methods
Electromagnetic field quantization
In this section, we briefly describe the quantization scheme of the electromagnetic field in the presence of arbitrary linear dielectric media, which follows that of ref. 27. The free field is characterized by a set of bosonic modes with frequencies ω and annihilation operators (j=x, y, z), with a corresponding Hamiltonian
The bosonic operators satisfy canonical commutation relations and are physically associated with noise polarization sources in the material. The electric field operator at frequency ω is driven by the sources at the same frequency and is given in the Coulomb gauge by^{27}
Here and in the following, it is assumed that all repeated vector or tensor indices are summed over, for example, is the frequencydependent permittivity at position r, and G_{ij} is the dyadic Green’s function which is the (gaugeinvariant) solution to
The total electric field is E(r)=∫dω E(r, ω)+h.c. Similarly, the magnetic field is . It can be verified that these expressions preserve the same field commutation relations as in freespace field quantization.
We now consider a twolevel atom at position r_{a} and with ground and excited states g›, e›, respectively, interacting with the electromagnetic field within the electric dipole approximation. The interaction Hamiltonian is given by
where is the atomic dipole matrix element.
Planewave expansion
We consider an interface with vacuum in the region z>0 and a dielectric with permittivity ε(ω) in the region z<0. The solution to equation (11) in the region z>0 can be written as the sum of a free component (that is, the solution in isotropic vacuum) and a scattered component. In the region 0≤z≤z′, they can be expanded in plane waves, which yields
Here k_{}=(k_{x}, k_{y}) and ρ=(x, y) are the wavevector and position along the direction parallel to the interface, while the full wavevector is given by with the relation . Further defining as the perpendicular component of the wavevector in the medium, and are the Fresnel reflection coefficients for s and p polarized waves, respectively.
Groundstate interactions
The interaction Hamiltonian H_{af} couples the ground state of the bare system g, 0› (consisting of the atomic ground state and vacuum) to states . Within secondorder perturbation theory, this interaction induces a shift of the bare groundstate energy by an amount
where ω_{0} is the bare resonance frequency of the atom. Substituting equations (10) and (13) into the expression above, and using analyticity of the integrand to rotate the frequency integral onto the positive imaginary axis ω=iu, one finds^{27}
Here, we have only included the scattered component of the Green’s function, as the free component gives a contribution that is independent of position. While we have considered a twolevel atom for simplicity, a more realistic model of an isotropic atom consists of multiple excited states and equal polarizabilities in each direction (that is, the ground state can emit virtual photons of any polarization with equal strength). Since the contribution to the groundstate shift from each excitedstate transition is additive, equation (17) can be modified to account for this by setting for all i and associating with the freespace spontaneous emission rate of any excited state. This yields equation (1).
Excitedstate interactions
We want to analyse the effective excitedstate properties in the specific case that an atom is weakly driven by a classical field of frequency ω_{L}. We thus transform to a rotating frame where the free atomic evolution is given by and δ=ω_{L}−ω_{0} is the detuning between the laser frequency and atomic resonance frequency ω_{0}.
The standard technique to derive the excitedstate shift and decay rate follows that of the ground state and uses timeindependent secondorder perturbation theory^{27}. In this approach, however, one cannot find the jump operator associated with photon emission, which is needed to calculate the effect of emission on atomic motion (for example, recoil heating). To rectify this, we employ an alternative approach based on deriving an atomic master equation within the Born–Markov approximation^{48}. In particular, the equation of motion for the reduced atomic density matrix ρ_{a} is given by
where is the interaction Hamiltonian in the interaction picture (with respect to the free Hamiltonian H_{0}=H_{a}+H_{f}) and 0› is the electromagnetic vacuum state. Any exponential of time appearing in the integral can be evaluated using the relation . Using the field expansion of equation (10) in and following some manipulation, equation (18) can be written in the form
The effective Hamiltonian describes coherent interactions between the atom and vacuum field and takes the form
which can be interpreted as a positiondependent energy shift of the excited state. Taking a simple isotropic twolevel model for an atom, the excitedstate shift is found to be
as stated in the main text.
The Liouvillian term [ρ_{a}] can be written in a Lindblad form
The jump operators take the form
The excitedstate spontaneous emission rate is in turn given through the relation , or
This again averages over dipole orientations and reproduces equation (3), with the replacement of ω_{0}→ω_{L}. To derive this, we have used the identity
The expressions for the excitedstate emission rate Γ(r_{a}) and level shift δω_{e}(r_{a}) agree with previous derivations^{27}, if the laser frequency ω_{L} is replaced by the atomic resonance frequency ω_{0}. The appearance of ω_{L} physically describes the effect of the weakly driven atom (that is, a linear optical element) interacting with its own Rayleighscattered field, which can be altered owing to the presence of nearby dielectric surfaces. This is in accordance with the wellknown result for the fluorescence spectrum of a driven atom, which is dominated by the Rayleigh contribution for atomic populations σ_{ee}<<1 (ref. 48). More importantly, we have obtained for the first time a decomposition of the jump operator associated with photon emission for arbitrary dielectric surroundings, which enables one to calculate recoil heating.
Shortdistance limit
We now derive the shortdistance expansion of the excitedstate properties. The resonant term in equations (2) and (3) depend on the quantity Tr G_{sc}(r_{a}, r_{a}, ω_{L}), which for planar interfaces is given by
For short distances, the expression is dominated by large k_{}≈ik_{⊥}, which gives rise to a long exponential tail . Expanding to second order for large k_{}, this tail contributes as
We now apply these results to the Drude model. First, we note the exact relation . For frequencies within a range Δ_{p}≲Q of the plasmon resonance, one readily finds that the response is well approximated by a complex Lorentzian. For larger detunings Δ_{p}≳Q, the exact expression of ε(ω) must be used. Furthermore, within the range Δ_{p}≲Q, one must compare the size of the first and second terms on the righthand side of equation (27). The first term dominates the integral provided that , in which case
which reproduces the asymptotic expressions for the excitedstate shift and linewidth in the main text.
Tunnelling
The atomic escape rate from the metastable potential formed by the vacuum trapping scheme owing to quantum tunnelling is calculated in the Wentzel–Kramers–Brillouin approximation in two steps. First, we find an approximation to the real part of the boundstate energy ħω_{b}. This is implemented by replacing the actual metastabledressedstate potential U_{d}(z) with a stable one, U_{s}(z), according to the formula (also see Fig. 4)
Here, z_{max} is the position corresponding to the maximum of the metastable barrier. The groundstate energy of U_{s}(z) is solved numerically.
Once the approximate groundstatebinding energy is obtained, the tunnelling probability P_{t} for a particle hitting the metastable barrier is calculated using the Wentzel–Kramers–Brillouin approximation,
Here, z_{1,2} are the solutions to U_{d}(z)=ħω_{b} and are thus the classical turning points of a particle with energy ħω_{b}. The rate that the trapped atom collides with the barrier is approximated by Δp_{z}/mΔz, where the position and momentum uncertainties are obtained from the numerical groundstate solution. We then estimate the inverse of the tunnellinglimited lifetime to be .
Recoil heating
Recoil heating can be calculated by treating the atomic position r_{a} in equation (23) as an operator. For notational simplicity, we shall consider recoil heating only along the trapping direction z, although our results can be easily generalized. The increase in momentum uncertainty owing to photon scattering is given by
Here we will focus solely on the Liouvillian term in the density matrix evolution (that is, momentum incurred from spontaneous emission). It can be shown that the Liouvillian term produces no net force, that is, . Using the fact that for any function f, one can derive the following general expression for the recoil heating rate in terms of Green’s functions,
Here, the derivatives ∂_{1} and ∂_{2} act on the first and second spatial arguments of the Green’s functions, respectively, and we have suppressed the atomic spatial variables in the directions that are not of interest. We further assume that the internal and spatial degrees of freedom can be decorrelated, . This is valid in our regime of interest where properties such as the excitedstate decay rate have negligible variation over the spatial extent of the atomic wave packets. As in the main text, we use the notation to denote the expectation value of atomic operators. Formally, we can write the above equation in the convenient form
This form is intuitive as it describes the momentum increase arising from a random walk process, where the atom experiences random momentum kicks of size ±ħk_{eff} at a rate Γ(r_{a})σ_{ee} (that is, the photonscattering rate). k_{eff} is a derived quantity that thus characterizes the effective momentum associated with scattered photons in the vicinity of a dielectric surface.
As a simple example, we consider an atom in free space polarized along x. Evaluating equation (33) with the freespace Green’s function G_{free,xx} from equation (14) yields (and Γ(r_{a})=Γ_{0}), recovering the known result^{43}. This recoil momentum is smaller than the full momentum ω_{L}/c of the scattered photon because the atom emits in a dipole pattern, but only the projection of the photon momentum along z contributes to motional heating in that direction. Applying the same formalism to an atom near a dielectric described by the Drude model, we find that the scattered component of the Green’s function dominates the recoil heating, leading to the asymptotic expansion of given in the main text. In all of our numerical calculations, the full Green’s function is used rather than asymptotic results.
Transient heating
Following a spontaneous emission event, the spatial wavefunction of the atom is temporarily untrapped and evolves under the pure groundstate potential for a characteristic time t_{trans} before the internal dynamics equilibrates. Here we derive an approximate propagator U(z_{f}, z_{i}, t_{f}, t_{i}) for the wavefunction evolving under the groundstate potential.
The problem significantly simplifies if the groundstate potential U_{g}(z)=ħω_{g}(z) is linearized around the atom trapping position z=z_{t},
This linearization is justified in our regime of interest where the atom is tightly trapped and the wavefunction undergoes little evolution in the time t_{trans}. The constant energy and position offsets in the Hamiltonian play no role, and thus we can consider the simplified Hamiltonian . The propagator for this Hamiltonian is given by
where is the wellknown propagator for a free particle^{49}.
In our simulations, the groundstate wavefunction for an arbitrary trapping potential is obtained numerically on a lattice, and so the propagator must be spatially discretized as well. In particular, while the exponential term in equation (36) is wellbehaved on a lattice, the free propagation U_{free} is not at short times. We instead replace U_{free} with the propagator D_{mn}(t) for the free discrete Schrodinger equation,
where β=ħ/2ma^{2}, n is the lattice site index, and a is the lattice constant. The propagator for this equation (defined as the matrix satisfying ψ_{m}(t)=D_{mn}ψ_{n}(0)) is given by
where J_{n} is the nth order Bessel function.
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How to cite this article: Chang, D. E. et al. Trapping atoms using nanoscale quantum vacuum forces. Nat. Commun. 5:4343 doi: 10.1038/ncomms5343 (2014).
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Acknowledgements
We thank N. Stern and O. Painter for helpful discussions. D.E.C. acknowledges support from Fundació Privada Cellex Barcelona. K.S. was funded by the NSF Physics Frontier Center at the JQI. J.M.T. acknowledges funding from the NSF Physics Frontier Center at the JQI and the US Army Research Office MURI award W911NF0910406. H.J.K. acknowledges funding from the IQIM, an NSF Physics Frontier Center with support of the Moore Foundation, by the AFOSR QuMPASS MURI, by the DoD NSSEFF program, and by NSF PHY1205729.
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D.E.C. and H.J.K. conceived of the idea. D.E.C. and K.S. performed the calculations. All authors contributed to the elaboration of the project and helped in editing the manuscript.
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Chang, D., Sinha, K., Taylor, J. et al. Trapping atoms using nanoscale quantum vacuum forces. Nat Commun 5, 4343 (2014). https://doi.org/10.1038/ncomms5343
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