Observation of a Discrete Time Crystal

Jan 30, 2017 | Publisher: Newswire Expert | Category: Science |  | Collection: Science News | Views: 14 | Likes: 2

Observation of a Discrete Time Crystal J. Zhang, P. W. Hess, A. Kyprianidis, P. Becker, A. Lee, J. Smith, G. Pagano,1 I.-D. Potirniche,2 A. C. Potter,2, 3 A. Vishwanath,2, 4 N. Y. Yao,2 and C. Monroe1 1Joint Quantum Institute, University of Maryland Department of Physics and National Institute of Standards and Technology, College Park, MD 20742 2Department of Physics, University of California Berkeley, Berkeley, CA 94720, USA 3Department of Physics, University of Texas at Austin, Austin, TX 78712, USA 4Department of Physics, Harvard University, Cambridge, MA 02138, USA (Dated: September 29, 2016) Spontaneous symmetry breaking is a funda- mental concept in many areas of physics, rang- ing from cosmology and particle physics to con- densed matter1. A prime example is the breaking of spatial translation symmetry, which underlies the formation of crystals and the phase transi- tion from liquid to solid. Analogous to crystals in space, the breaking of translation symmetry in time and the emergence of a “time crystal” was recently proposed2,3, but later shown to be for- bidden in thermal equilibrium4–6. However, non- equilibrium Floquet systems subject to a periodic drive can exhibit persistent time-correlations at an emergent sub-harmonic frequency7–10. This new phase of matter has been dubbed a “discrete time crystal” (DTC)10,11. Here, we present the first experimental observation of a discrete time crystal, in an interacting spin chain of trapped atomic ions. We apply a periodic Hamiltonian to the system under many-body localization (MBL) conditions, and observe a sub-harmonic temporal response that is robust to external perturbations. Such a time crystal opens the door for studying systems with long-range spatial-temporal corre- lations and novel phases of matter that emerge under intrinsically non-equilibrium conditions7. For any symmetry in a Hamiltonian system, its spon- taneous breaking in the ground state leads to a phase transition12. The broken symmetry itself can assume many different forms. For example, the breaking of spin- rotational symmetry leads to a phase transition from paramagnetism to ferromagnetism when the temperature is brought below the Curie point. The breaking of spatial symmetry leads to the formation of crystals, where the continuous translation symmetry of space is replaced by a discrete one. We now pose an analogous question: can the trans- lation symmetry of time be broken? The proposal of such a “time crystal”2 for time-independent Hamiltoni- ans has led to much discussion, with the conclusion that such structures cannot exist in the ground state or any thermal equilibrium state of a quantum mechanical sys- tem4–6. A simple intuitive explanation is that quantum equilibrium states have time-independent observables by Tim e (100 Floquet periods) H1 H2 H3 H1 = g(1 - ε) σi y H2 = Jij σi x σj x H3 = Di σi x H1 H2 H3 H1 H2 H3 Global rotation Interactions Disorder FIG. 1: Floquet evolution of a spin chain. Three Hamil- tonians are applied sequentially in time: a global spin flip of nearly pi (H1), long-range Ising interactions (H2), and strong disorder (H3). The system evolves for 100 Floquet periods of this sequence. construction; thus, time translation symmetry can only be spontaneously broken in non-equilibrium systems7–10. In particular, the dynamics of periodically-driven Flo- quet systems possesses a discrete time translation sym- metry governed by the drive period. This symmetry can be further broken into “super-lattice” structures where physical observables exhibit a period larger than that of the drive. Such a response is analogous to commensurate charge density waves that break the discrete translation symmetry of their underlying lattice1. The robust sub- harmonic synchronization of the many-body Floquet sys- tem is the essence of the discrete time crystal phase7–10. In a DTC, the underlying Floquet drive should gener- ally be accompanied by strong disorder, leading to many- body localization13 and thereby preventing the quantum system from absorbing the drive energy and heating to infinite temperatures14–17. Here, we report the direct observation of discrete time translation symmetry breaking and DTC formation in a spin chain of trapped atomic ions, under the influence of a periodic Floquet-MBL Hamiltonian. We experimen- tally implement a quantum many-body Hamiltonian with long-range Ising interactions and disordered local effec- tive fields, using optical control techniques19,20. Follow- ar X iv :1 60 9. 08 68 4v 1 [q ua nt- ph ] 27 Se p 2 01 6 2 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 (a) W t3 = π ε = 0.03, 2πJ0 t2 /(W t3)=0.07 (b) (c) (d) (e) ε = 0.11, 2πJ0 t2 /(W t3)=0.07 Ion # 1 Ion # 2 Ion # 3 Ion # 4 Ion # 5 Ion # 6 Ion # 7 Ion # 8 Ion # 9 Ion # 10 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 Time (T) S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 S in gl e Io n M ag ne tiz at io ns 0.4 0.5 0.6 0 0.02 0.04 0.06 Frequency 0.4 0.5 0.6 0 0.02 0.04 0.06 Frequency (1/T) 0.4 0.5 0.6 0 0.02 0.04 0.06 Frequency 0.4 0.5 0.6 0 0.02 0.04 0.06 Interactions On (1/T) Frequency (1/T) (1/T) Time (T) Time (T) Time (T) Ion # 3 Ion # 8 (f) 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 ε = 0.03, W t3 = 0 ε = 0.03, FF T sp ec tru m FF T sp ec tru m FF T sp ec tru m FF T sp ec tru m Interactions Off Time (T) Time (T) Time (T) Time (T) Time (T) Time (T) M ag ne tiz at io n M ag ne tiz at io n M ag ne tiz at io n M ag ne tiz at io n FIG. 2: Spontaneous breaking of discrete time translation symmetry. Top panel: Time-evolved magnetizations of each spin 〈σxi (t)〉 and their Fourier spectra, showing sub-harmonic response of the system to the Floquet Hamiltonian. (a) When only the H1 spin flip is applied, the spins oscillate with a sub-harmonic response that beats due to the perturbation ε = 0.03 from perfect pi-pulses, with a clear splitting in the Fourier spectrum. (b) With both the H1 spin flip and the disorder H3, the spins precess with various Larmor rates in the presence of different individual fields. (c) Finally, adding the spin-spin interaction term H2 (shown with the largest interaction phase J0t2 = 0.036 rad), the spins lock to the sub-harmonic frequency of the drive period. Here the Fourier spectrum merges into a single peak even in the face of perturbation ε on the spin flip H1. (d) When the perturbation is too strong (ε = 0.11), we cross the boundary from the discrete time crystal into a symmetry unbroken phase10. Bottom panel: Individually resolved time traces. (e) Spin magnetization for all 10 spins corresponding to the case of (b). (f) Spin 3 and 8 corresponding to the case of (c). Each point is the average of 150 experimental repetitions. Error bars are computed from quantum projection noise and detection infidelities. ing the evolution through many Floquet periods, we mea- sure the temporal correlations of the spin magnetization dynamics. A DTC requires the ability to control the interplay be- tween three key ingredients: strong drive, interactions, and disorder. These are reflected in the applied Floquet Hamiltonian H, consisting of the following three succes- sive pieces with overall period T = t1 + t2 + t3 (see Fig. 1) (~ = 1): H =  H1 = g(1− ε) ∑ i σ y i , time t1 H2 = ∑ i Jijσ x i σ x j , time t2 H3 = ∑ iDiσ x i time t3. (1) Here, σγi (γ = x, y, z) is the Pauli matrix acting on the ith spin, g is the Rabi frequency with small perturbation ε, Jij is the coupling strength between spins i and j, and Di is a site-dependent disordered potential sampled from a uniform random distribution with Di ∈ [0,W ]. To implement the Floquet Hamiltonian, each of the effective spin-1/2 particles in the chain is encoded in the 2S1/2 |F = 0,mF = 0〉 and |F = 1,mF = 0〉 hyperfine ‘clock’ states of a 171Yb+ ion, denoted |↓〉z and |↑〉z and separated by 12.642831 GHz21. We store a chain of 10 ions in a linear rf Paul trap, and apply single spin rota- tions using optically-driven Raman transitions between the two spin states22. Spin-spin interactions are gener- ated by spin-dependent optical dipole forces, which give rise to a tunable long-range Ising coupling23,24 that falls off approximately algebraically as Jij ∝ J0/|i− j|α. Pro- grammable disorder among the spins is generated by the ac Stark shift from a tightly focused laser beam that ad- dresses each spin individually20. The Stark shift is an effective site-dependent σzi field, so we surround this op- eration with pi2 -pulses to transform the field into the x direction of the Bloch sphere (see Methods). Finally, we measure the magnetization of each spin by collecting the spin-dependent fluorescence on a camera for site-resolved imaging. This allows access to the single-site magnetiza- 3 tion, σγi , along any direction with a detection fidelity > 98% per spin. The unitary time evolution under a single Floquet pe- riod is U(T ) = e−iH3t3e−iH2t2e−iH1t1 . (2) The first evolution operator e−iH1t1 nominally flips all the spins around the y-axis of the Bloch sphere by an angle 2gt1 = pi, but also includes a controlled perturba- tion in the angle, εpi, where ε < 0.15. This critical ro- tation step is susceptible to noise in the Rabi frequency (1% rms) from laser intensity instability, and also op- tical inhomogeneities (< 5%) across the chain due to the shape of the Raman laser beams. In order to ac- curately control H1, we use the BB1 dynamical decou- pling echo sequence25 (see Methods) to suppress these effects, resulting in control of the rotation angle to a pre- cision < 0.5%. The second evolution operator e−iH2t2 applies the spin-spin Ising interaction, where the maxi- mum nearest-neighbor coupling J0 ranges from 2pi(0.04 kHz) to 2pi(0.25 kHz) and decays with distance at a power law exponent α = 1.51. The duration of the interaction term is set so that J0t2 < 0.04 rad of phase accumulation. The third evolution operator e−iH3t3 provides disorder to localize the system, and is programmed so that the vari- ance of the disorder is set by Wt3 = pi. In this regime, MBL in the thermodynamic limit is expected to persist even in the presence of long-range interactions26–28. To observe the DTC, we initialize the spins to the state |ψ0〉 = |↓〉x = 1√2 (|↓〉z + |↑〉z) through optical pumping followed by a global pi2 rotation. After many periods of the above Floquet unitary Eq.(2), we measure the mag- netization of each spin along x, which gives the time- correlation function 〈σxi (t)〉 = 〈ψ0|σxi (t)σxi (0) |ψ0〉 . (3) Figure 2 depicts the measured spin magnetization dy- namics, both in the time and the frequency domain, up to N = 100 Floquet periods. A single Floquet period T is set to a value between 74-75 µs, depending on the parameters in the Hamiltonian. The global pi-pulse e−iH1t1 rotates the spins roughly half way around the Bloch sphere, so that we expect a response of the system at twice the drive period 2T , or half of the Floquet frequency. The frequency of this sub- harmonic response in the magnetization is sensitive to the precise value of the global rotation in H1 and is therefore expected to track the perturbation ε. This results in co- herent beats and a splitting in the Fourier spectrum by 2ε (Fig. 2(a)). When we add disorder e−iH3t3 to the Flo- quet period, the single spins precess at different Larmor rates (Fig. 2(e)) and dephase with respect to each other (Fig. 2(b)). Only upon adding Ising interactions e−iH2t2 , and hence many-body correlations, the spin synchroniza- tion is restored (Fig. 2(c,f)). 0 0.05 0.1 0.00 0.02 0.04 0.06 0.08 0 0.05 0.1 0.00 0.02 0.04 0.06 0.08 0 0.05 0.1 0.00 0.02 0.04 0.06 0.08 0 0.05 0.1 0.00 0.02 0.04 0.06 0.08 0.048 0.012 0.072 2πJ0 t2 /(W t3) 2πJ0 t2 /(W t3) 2πJ0 t2 /(W t3) 2πJ0 t2 /(W t3) ε ε ε ε 0.024 (a) P ea k he ig ht va ria nc e (a rb .) (b) 2πJ0 t2 /(W t3) 0.05 0.1 0 0.02 0.04 0.06 P er tu rb at io n ε Thermal Discrete Time Crystal Symmetry Unbroken MBL FIG. 3: Variance of the subharmonic peak amplitude as a signature of the DTC transition. (a) Variances of the central peak amplitude, computed over the 10 sites and averaged over 10 instances of disorder, for four different strengths of the long-range interation term J0. The cross- over from symmetry unbroken state to a DTC is observed as a peak in the measured variance of the sub-harmonic sys- tem response. Dashed lines: numerical results, scaled verti- cally to fit the experimental data (see supplementary infor- mation). Experimental error bars are standard error of the mean. (b) Cross-over determined by a fit to the variance peak location (dots). Dashed lines: numerically determined phase boundaries with experimental long-range coupling parame- ters10. Grey shaded region indicates 90% confidence level of the DTC to symmetry unbroken phase boundary. The key result is that with all of these elements, the temporal response is locked to twice the Floquet period, even in the face of perturbations to the drive in H1. This can be seen clearly as the split Fourier peaks from Fig. 2(b) merge into a single peak in Fig. 2(c). This repre- sents the “rigidity” of the discrete time crystal10, which persists under moderate perturbation strengths. How- ever, for large ε, the DTC phase disappears as evinced by the decay of the sub-harmonic temporal correlations and the suppression of the central peak heights, as shown in Fig. 2(d). In the thermodynamic limit, these pertur- bations induce a phase transition from a DTC to a sym- metry unbroken MBL phase7–10, which is rounded into a cross-over in finite size systems. The phase boundary is defined by the competition be- tween the drive perturbation ε and strength of the inter- actions J0. We probe this boundary by measuring the variance of the sub-harmonic spectral peak height, com- puted over the 10 sites and averaged over 10 instances of disorder. Figure 3(a) shows the variances as a func- tion of the perturbation ε, for four different interaction 4 strengths. As we increase ε, the variance growth distinc- tively captures the onset of the transition, with increased fluctuations signaling the crossing of the phase bound- ary. When the perturbations are too large, the crystal “melts”. Figure 3(b) shows the fitted centers of the vari- ance curve, on top of numerically computed phase bound- aries with experimental parameters. The measurements are in agreement with the expected DTC to time crystal “melting” boundary, which displays approximately linear dependence on the perturbation strength in the limit of small interactions10. 0.00 0.02 0.04 0.06 0.08 0.10 0.12 0.00 0.02 0.04 0.06 0.08 P ea k he ig ht (a rb .u ni ts ) Perturbation ε 2πJ0 t2 /(W t3) 0.072 0.048 0.024 0.012 FIG. 4: Subharmonic peak amplitude as a function of the drive perturbation. Central subharmonic peak ampli- tude in the Fourier spectrum as a function of the perturba- tion ε, averaged over the 10 sites and 10 disorder instances, for four different interaction strengths. Solid lines are guides to the eye. The height decreases across the phase boundary and eventually diminishes as the single peak is split into two. Error bars: 1 s.d. Figure 4 illustrates the amplitude of the subharmonic peak as a function of ε, for the four different applied in- teraction strengths. The amplitude drops off with larger perturbations, and this slope is steeper as we decrease the interaction strength. The sub-harmonic peak ampli- tude observable in Fig. 4 can serve as an auxilary probe of the rigidity. It is expected to be an order parameter for the DTC phase and thus, to scale similarly to the mutual information between spins8,10. This connection also provides insight into the Floquet many-body quan- tum dynamics, and in particular to the correlations or entanglement that underly the DTC phase. Indeed, the eigenstates of the entire Floquet unitary are expected to resemble GHZ or spin-“Schro¨dinger Cat” states8,18. The initial product state in the experiment can be written as a superposition of two cat states: |↓↓ · · · ↓〉x = 1√2 (|φ+〉+ |φ−〉), where |φ±〉 = 1√2 (|↓↓ · · · ↓〉x ± |↑↑ · · · ↑〉x). These two states evolve at different rates corresponding to their respective quasi-energies, giving rise to the sub-harmonic periodic oscillations of physical observables. Such oscil- lations are expected to persist at increasingly long times as the system size increases7,8,10. In summary, we present the first experimental obser- vation of discrete time translation symmetry breaking into a DTC. We measure persistent oscillations and syn- chronizations of interacting spins in a chain and show that the discrete time crystal is rigid, or robust to perturbations in the drive. Our Floquet-MBL system with long-range interactions provides an ideal testbed for out-of-equilibrium quantum dynamics and the study of novel phases of matter that exist only in a Floquet set- ting7–10,30,31. Such phases can also exhibit topological order31–35 and can be used for various quantum infor- mation tasks, such as implementing a robust quantum memory36. ACKNOWLEDGEMENTS We acknowledge useful discussions with Mike Zaletel and Dan Stamper-Kurn. This work is supported by the ARO Atomic and Molecular Physics Program, the AFOSR MURI on Quantum Measurement and Verifica- tion, the IARPA LogiQ program, the IC Postdoctoral Research Fellowship Program, the NSF Physics Frontier Center at JQI, and the Miller Institute for Basic Re- search in Science. A. V. was supported by the AFOSR MURI grant FA9550- 14-1-0035 and Simons Investigator Program. [1] Chaikin, P. & Lubensky, T. Principles of Condensed Matter Physics, vol. 1 (Cambridge University Press, 1995). [2] Wilczek, F. Quantum time crystals. Phys. Rev. Lett. 109, 160401 (2012). [3] Wilczek, F. Superfluidity and space-time translation symmetry breaking. Phys. Rev. Lett. 111, 250402 (2013). [4] Bruno, P. Comment on “quantum time crystals”. Phys. Rev. Lett. 110, 118901 (2013). [5] Bruno, P. Impossibility of spontaneously rotating time crystals: A no-go theorem. Phys. Rev. Lett. 111, 070402 (2013). [6] Watanabe, H. & Oshikawa, M. Absence of quantum time crystals. Phys. Rev. Lett. 114, 251603 (2015). [7] Khemani, V., Lazarides, A., Moessner, R. & Sondhi, S. L. Phase structure of driven quantum systems. Phys. Rev. Lett. 116, 250401 (2016). [8] Else, D. V., Bauer, B. & Nayak, C. Floquet time crystals. Phys. Rev. Lett. 117, 090402 (2016). [9] von Keyserlingk, C. W., Khemani, V. & Sondhi, S. L. Absolute stability and spatiotemporal long-range order in floquet systems. Phys. Rev. B 94, 085112 (2016). [10] Yao, N. Y., Potter, A. C., Potirniche, I.-D. & Vish- wanath, A. Discrete time crystals: rigidity, criticality, and realizations. arXiv:1608.02589 (2016). [11] This phase is also referred to as a pi-spin glass7 or a Flo- quet time crystal8. 5 [12] Sachdev, S. Quantum Phase Transitions (Cambridge Univ. Press, Cambridge, 1999). [13] Nandkishore, R. & Huse, D. A. Many-body localiza- tion and thermalization in quantum statistical mechan- ics. Annual Review of Condensed Matter Physics 6, 15– 38 (2015). [14] D’Alessio, L. & Rigol, M. Long-time behavior of isolated periodically driven interacting lattice systems. Phys. Rev. X 4, 041048 (2014). [15] Ponte, P., Chandran, A., Papic, Z. & Abanin, D. A. Pe- riodically driven ergodic and many-body localized quan- tum systems. Annals of Physics 353, 196 – 204 (2015). [16] Eckardt, A. Atomic quantum gases in periodically driven optical lattices. arXiv:1606.08041v1 (2016). [17] We note that under certain conditions, time-crystal dy- namics can persist for rather long times even in the ab- sence of localization before ultimately being destroyed by thermalization18. [18] Else, D. V., Bauer, B. & Nayak, C. Pre-thermal time crystals and floquet topological phases without disorder. arXiv:1607.05277 (2016). [19] Smith, J. et al. Many-body localization in a quantum simulator with programmable random disorder. Nat. Phys. advance online publication, – (2016). [20] Lee, A. C. et al. Engineering large stark shifts for con- trol of individual clock state qubits. arXiv:1604.08840 (2016). [21] Olmschenk, S. et al. Manipulation and detection of a trapped Yb+ hyperfine qubit. Phys. Rev. A 76, 052314 (2007). [22] Hayes, D. et al. Entanglement of atomic qubits using an optical frequency comb. Phys. Rev. Lett. 104, 140501 (2010). [23] Porras, D. & Cirac, J. I. Effective quantum spin systems with trapped ions. Phys. Rev. Lett. 92, 207901 (2004). [24] Kim, K. et al. Entanglement and tunable spin-spin cou- plings between trapped ions using multiple transverse modes. Phys. Rev. Lett. 103, 120502 (2009). [25] Brown, K. R., Harrow, A. W. & Chuang, I. L. Arbitrarily accurate composite pulse sequences. Phys. Rev. A 70, 052318 (2004). [26] Burin, A. L. Localization in a random xy model with long-range interactions: Intermediate case between single-particle and many-body problems. Phys. Rev. B 92, 104428 (2015). [27] Yao, N. Y. et al. Many-body localization in dipolar sys- tems. Phys. Rev. Lett. 113, 243002 (2014). [28] There has been recent work29 questioning the stability of MBL with anything longer-range than exponential-in- distance interaction. However, the proposed mechanism is not relevant on experimentally accessible length- and time- scales. [29] De Roeck, W. & Huveneers, F. Stability and instability towards delocalization in mbl systems. arXiv:1608.01815 (2016). [30] Bordia, P. & Luschen, H. & Schneider, U. & Kanp, M. & Bloch, I. Periodically Driving a Many-Body Localized Quantum System. arXiv:1607.07868 (2016). [31] von Keyserlingk, C. W. & Sondhi, S. L. Phase structure of one-dimensional interacting floquet systems. i. abelian symmetry-protected topological phases. Phys. Rev. B 93, 245145 (2016). [32] Else, D. V. & Nayak, C. Classification of topolog- ical phases in periodically driven interacting systems. arXiv:1602.04804 (2016). [33] von Keyserlingk, C. W. & Sondhi, S. L. Phase struc- ture of one-dimensional interacting floquet systems. ii. symmetry-broken phases. Phys. Rev. B 93, 245146 (2016). [34] Potter, A. C. & Morimoto, T. & Vishwanath, A. Topo- logical classification of interacting 1D Floquet phases. arXiv:1602.05194 (2016). [35] Roy, R. & Harper, F. Abelian Floquet symmetry- protected topological phases in one dimension. Phys. Rev. B 94, 125105 (2016). [36] Bahri, Y. & Vosk, R. & Altman, E. & Vishwanath, A. Localization and topology protected quantum coherence at the edge of hot matter. Nature Communications 6, 7341 (2016). METHODS SUMMARY Dynamical decoupling sequence. We use a pair of Raman laser beams globally illuminat- ing the entire ion chain to drive qubit rotations22. The ion chain has 25 µm length, and we shape the beams to have 200 µm full width half maximum along the ion chain, resulting in ∼ 5% intensity inhomogeneity. When a fixed duration is set for H1 in Eq. 1 of the main text, the time dependent magnetization for different ions ac- quire different evolution frequencies, resulting in the net magnetization of the system dephasing after about 10 pi-pulses. In addition, the Raman laser has rms inten- sity noise of about 1%, which restricts the spin-rotation coherence to only about 30 pi-pulses (80% contrast). To mitigate these imperfections, we employ a BB1 dy- namical decoupling pulse sequence for the drive unitary U1 (written for each spin i): U1(ε) = e −iH1t1 = e−i pi 2 σ θ i e−ipiσ 3θ i e−i pi 2 σ θ i e−i pi 2 (1−ε)σyi , where in addition to the perturbed pi rotation e−i pi 2 (1−ε)σyi , three additional rotations are applied: a pi- pulse along an angle θ = arccos(−pi 4pi ), a 2pi-pulse along 3θ, and another pi-pulse along θ, where the axes of these additional rotations are in the x-y plane of the Bloch sphere with the specified angle referenced to the x-axis. In this way, any deviation in the original rotation from the desired value of pi(1− ε) is reduced to second order. See Ref.25 for detailed discussions. Generating the effective Ising Hamiltonian We generate spin-spin interactions by applying spin- dependent optical dipole forces to ions confined in a 3- layer linear Paul trap with a 4.8 MHz transverse center- of-mass motional frequency. Two off-resonant laser beams with a wavevector difference ∆~k along a princi- pal axis of transverse motion globally address the ions 6 and drive stimulated Raman transitions. The two beams contain a pair of beatnote frequencies symmetrically de- tuned from the spin transition frequency by an amount µ, comparable to the transverse motional mode frequencies. In the Lamb-Dicke regime, this results in the Ising-type Hamiltonian in Eq. (1)23,24 with Ji,j = Ω 2ωR N∑ m=1 bi,mbj,m µ2 − ω2m , (4) where Ω is the global Rabi frequency, ωR = ~∆k2/(2M) is the recoil frequency, bi,m is the normal-mode matrix, and ωm are the transverse mode frequencies. The cou- pling profile may be approximated as a power-law decay Ji,j ≈ J0/|i − j|α, where in principle α can be tuned between 0 and 3 by varying the laser detuning µ or the trap frequencies ωm. In this work, α is fixed at 1.51 by setting the axial trapping frequency to be 0.44 MHz, and Raman beatnode detuning to be 155 kHz. Apply disorder in the axial direction. We apply the strong random disordered field with a fourth-order ac Stark shift20, which is naturally an effec- tive σzi operator. To transform this into a σ x i operator, we apply additional pi/2 rotations. Hence the third term in the Floquet evolution U3 (written for each spin i) is also a composite sequence: U3 = e −iH3t3 = ei pi 4 σ y i e−iDiσ z i t3e−i pi 4 σ y i = e−iDiσ x i t3 . METHODS Experimental time sequence. The Floquet time evolution is realized using the timing sequence illustrated in Fig. 5. The chain of 10 trapped ions is initialized in the ground motional state of their center of mass motion us- ing doppler cooling and sideband cooling (not shown). Optical pumping prepares the ions in the |↓〉z state. We then globally rotate each spin vector onto |↓〉x by per- forming a pi2 -pulse around the y-axis (“Initialization” in Fig. 5). The H1 Hamiltonian (perturbed pi-pulses) lasts 14-15 µs depending on ε and it consists of a four pulse BB1 sequence as described in the methods summary. Our carrier Rabi frequency is set such that we perform a pi- pulse in less than 3 µs. Including the three compensating pulses, the BB1 sequence requires 5 pi-pulse times to im- plement U1(ε). The H2 Ising Hamiltonian is applied for 25 µs (“Spin- spin interactions” in Fig. 5). The pulse time was suffi- ciently long that with the pulse shaping described below the effects of the finite pulse time spectral broadening were largely reduced. To compensate for the residual off-resonant carrier drive, we apply a small amplitude transverse field (“Compensation” in Fig. 5) for 2 µs. For the H3 disorder Hamiltonian, we apply ∑N i=1Diσ z i (“Strong random disorder” in Fig. 5) generated by Stark shifts as described in the Methods summary. This is sandwiched in-between rotations around the y-axis to convert this into disorder in σxi . After up-to 100 applica- tions of the Floquet evolution, we rotate the state back to the z-axis (“Prepare for measurement” in Fig. 5) and detect the spin state |↑〉z or |↓〉z using spin-dependent fluorescence. Pulse shaping for suppressing off-resonant excitations. The optical control fields for generating H1, H2 and H3 are amplitude modulated using acousto-optic modulators (AOMs) to generate the evolution operators. If the RF drive to these modulators is applied as a square pulse will be significantly broadened in the Fourier domain due to fast rise and fall times at the edges (100 ns). As the pulse duration decreases, the width of the spectral broadening will expand (see Fig 5 inset). The components of the evolution operator must be sufficiently short in order to evolve 100 Floquet periods within a decoherence time of < 8 ms. This spectral broadening is problematic when generat- ing the interaction Hamiltonian H2 due to off-resonant driving. The spin-spin interactions are applied using beatnote frequencies detuned 4.8 MHz from the carrier transition and 155 kHz from the sidebands. A broad pulse in the frequency domain can drive either the qubit hyperfine transition at the carrier frequency or phonons via sideband transitions. Similar issues occur while we apply the disordered field in H3. The fourth-order ac Stark shift is generated from a frequency comb that has a closest beatnote which is 23 MHz away from hyperfine and Zeeman transitions. To apply large average Stark shifts (33 kHz max.) across the 10 ions, we raster the laser beam once within a single cycle (30 µs), for a single-ion pulse duration of 3µs. The fast rastering also produces off-resonant carrier driving that resemble σyi fields. We mitigate both these effects by shaping the pulses with a 25% “Tukey” window , a cosine tapered function for the rise and fall. This largely removes off-resonant terms in all Hamiltonian chapters, while minimizing the reduction in pulse area (80%). We carefully character- ize any residual effects in H2 with a single ion where no interaction dynamics are present, and apply a small com- pensation field in σyi to cancel residual effects. We deduce an upper limit of 0.3% (relative to H1) on the residuals from the envelopes of the dynamically decoupled pi pulse 7 sequence. Emergence of time crystal stabilized by the interactions and the disorder. In the main text we have shown different scenarios which occur when we modify sub-sections of the total Floquet evolution operator. Here we first expand the results shown in Fig.2 of the main text by highlighting the effect of the interactions. Figure 6 shows individual magnetization evolutions with the same ε and instance of disorder {Di} but with increasing interaction J0, from left to right. This shows how, all else being equal, as we turn up the interaction strength the synchronization across the ions gradually builds up signaling the forma- tion of a discrete time crystal. In a complementary fashion, Fig. 7 addresses the role of disorder in stabilizing the time crystal formation. In particular we compare the magnetization dynamics, both in the time and frequency domain, including or removing the disorder chapter from the total Floquet evolution. We show on the side the corresponding exact numerics calculated by applying U(T )N on the initial state |ψ0〉 with the measured experimental parameters. Figure 7(a) shows that, with no disorder, the ions coherent dynamics tracks the perturbation ε, which results in coherent beats in agreement with our exact results. On the other hand, Figure 7(b) shows that, all else being equal, adding the disorder chapter locks the subharmonic response of all the ions. Although the numerics is qualitatively in agreement with the experimental data, nevertheless we observe a decay in the magnetization that cannot be explained by our numerical unitary simulations. This damping can be due to two possible sources: one is the residual off-resonant drive of the disorder chap- ter which is not totally eliminated by the pulse shap- ing (see pulse-shaping section above). This small resid- ual effect can behave like residual σxi and σ y i terms, or coupling between the clock spin states and the Zeeman states |F = 1,mF = ±1〉, resulting in a decay in the co- herent oscillations. This effect varies across the different disorder instances and the different interaction strengths and leads to an overall decrease of the Fourier subhar- monic peak height and of its variance with respect to what is expected from the theory (see Fig.7(b)). To take this effect into account in the theory, we perform a least squares fit of the amplitude of the four theory curves to the experimental data. With this procedure we obtain the scaling factors (0.56, 0.53, 0.51, 0.78) for the inter- action strengths J0t2/(Wt3) = (0.72, 0.48, 0.24, 0.12) re- spectively, which are consistent with the decays discussed above. Data analysis and fitting procedure The peak height variance data in Fig. 3(a) are fit to a lineshape in order to extract the cross-over transition boundary εp in Fig. 3(b). We use the following phe- nomenologial lineshape, a Lorentzian of the log10() F (ε) = A 1 1 + ( log10(ε/εp) γ )2 +B (5) A statistical error is extracted from a weighted non- linear fit to the data, which yields a fractional standard error bar of a few percent. The error in the peak height is limited by systematic error in the finite number of in- stances we realized in the experiment. For each value of J0 and ε we average over the same 10 instance of disorder. These 10 instances are averaged and fit to F (ε). In order to estimate the error due to the finite disorder instances, we perform random sampling from a numerical dataset of 100 instances of disorder. Sampling 10 of these and fitting the peak height variance data to F (ε) yields a Gaussian distribution of extracted peak centers p over 10000 repetitions (Fig. 8). We take the one standard deviation (67%) confidence interval in the sample as the systematic uncertainty in the fit, which yield a fractional error of ∼ 15%. This systematic uncertainty dominates over the statistical, so we apply error bars in Fig. 3(b) equal to this computed finite sampling error. 8 U1 U2 Spin-spin interactions Programmable random disorder Compensation Rotate z into x π 2π π π(1-ε) rotation Rotate x back to z Raster across 10 sites Pulse shaping U3 14~15 μs 27 μs 33 μs BB1 dynamical decoupling Time U (T) Single period Initialization Prepare for measurement Spin state detection (1 ms) optical pumping (15 μs) 101 Floquet periods Fourier spectrum Unshaped Shaped (Tukey window) FIG. 5: Experimental pulse sequence. See text for detailed explanations. 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 S in gl e Io n M ag ne tiz at io ns 0.4 0.5 0.6 0 0.02 0.04 0.06 S pe ct ra lD en si ty (a rb .) 0.4 0.5 0.6 0 0.02 0.04 0.06 S pe ct ra lD en si ty (a rb .) 0.4 0.5 0.6 0 0.02 0.04 0.06 S pe ct ra lD en si ty (a rb .) Frequency (1/T) Frequency (1/T) Time (T) Time (T) 2πJ0 t2 /(W t3)=0.024 2πJ0 t2 /(W t3)=0.07 ε = 0.035, 2πJ0 t2 /(W t3)=0.012 (a) (b) (c) Frequency (1/T) Time (T) ε = 0.035, ε = 0.035, FIG. 6: Build up of a discrete time-crystal. From (a) to (b) and (c), we fix the disorder instance and the perturbation, while gradually increase the interactions. The temporal oscillations are synchronized with increasing interactions, and the Fourier sub-harmonic peak is enhanced. 9 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 Time (T) S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 Time (T) S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 Time (T) S in gl e Io n M ag ne tiz at io ns 0 20 40 60 80 100 - 1.0 - 0.5 0.0 0.5 1.0 Time (T) S in gl e Io n M ag ne tiz at io ns 0.4 0.5 0.6 0 0.05 0.1 Frequency (1/T) S pe ct ra lD en si ty (a rb .) 0.4 0.5 0.6 0 0.05 0.1 Frequency (1/T) S pe ct ra lD en si ty (a rb .) 0.4 0.5 0.6 0 0.05 0.1 Frequency (1/T) S pe ct ra lD en si ty (a rb .) 0.4 0.5 0.6 0 0.05 0.1 Frequency (1/T) S pe ct ra lD en si ty (a rb .) Without Disorder With Disorder Experiment Exact numerics Experiment Exact numerics ε = 0.035, 2J0 t2 = 0.036, W t3=0 ε = 0.035, 2J0 t2 = 0.036, W t3=0 ε = 0.035, 2J0 t2 = 0.036, W t3=π ε = 0.035, 2J0 t2 = 0.036, W t3=π Ion 1 Ion 2 Ion 3 Ion 4 Ion 5 Ion 6 Ion 7 Ion 8 Ion 9 Ion 10 Ion 1 Ion 2 Ion 3 Ion 4 Ion 5 Ion 6 Ion 7 Ion 8 Ion 9 Ion 10 FIG. 7: Comparing ions dynamics with and without disorder. (a) Without disorder, the interaction suppresses the beating. Left: experimental data; right: exact numerics calculated under the Floquet time-evolution. (b) With disorder, the time-crystal is more stable. Left: experimnetal result for one single instance; right: numerical simulations. 0.00 0.05 0.10 0.15 0.20 0.25 0.30 0.00 0.01 0.02 0.03 0.04 Perturbation ε P ea k he ig ht va ria nc e (a rb .) 0.030 0.035 0.040 0.045 0.050 0.055 0.060 0 10 20 30 40 50 60 Fit Center (εp) P D F FIG. 8: A random sampling from numerical evolutions. Averages of 10 disorder instances from numerical evolution under H for 2piJ0t2/(Wt3) = 0.072. Left: An example random numerical dataset (points) and fit to Methods Eq. 5 (dashed line). Right: The normalized probability distribution (PDF) of peak fit centers εp is shown in yellow, and a normal distribution is overlaid in red. The normal distribution is fit using only the mean and standard deviation of the sample, showing excellent agreement with Gaussian statistics. For this value of J0 the mean ε¯p = 0.046 and the standard deviation σεp = 0.006.

Scientists have confirmed a brand new form of matter: time crystals

Perpetual motion without energy.

Observation of a Discrete Time Crystal

J. Zhang, P. W. Hess, A. Kyprianidis, P. Becker, A. Lee, J. Smith, G. Pagano,1 I.-D. Potirniche, 2 A. C. Potter, 2, 3 A. Vishwanath, 2, 4 N. Y. Yao, 2 and C. Monroe 1

1 Joint Quantum Institute, University of Maryland Department of Physics and National Institute of Standards and Technology, College Park, MD 20742
2 Department of Physics, University of California Berkeley, Berkeley, CA 94720, USA
3 Department of Physics, University of Texas at Austin, Austin, TX 78712, USA
4 Department of Physics, Harvard University, Cambridge, MA 02138, USA
(Dated: September 29, 2016)

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